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\begin{document}

\title{Incommensurate Twisted Bilayer Graphene: emerging quasi-periodicity and stability}
 
\author{Ian Jauslin}
\affiliation{Rutgers University, Department of Mathematics, New Brunswick, USA}
\email{ian.jauslin@rutgers.edu}

\author{Vieri Mastropietro}
\affiliation{Università di Roma ``La Sapienza'', Department of Physics, Rome, Italy }
\email{vieri.mastropietro@uniroma1.it}

\begin{abstract} We consider a lattice model of Twisted Bilayer Graphene (TBG). The presence of incommensurate angles produces
an emerging quasi-periodicity manifesting itself in large momenta Umklapp interactions that almost connect
the Dirac points. We rigorously establish the stability of the semimetallic phase
via a Renormalization Group analysis combined with number theoretical properties of irrationals, similar to the ones used in Kolmogorov-Arnold-Moser (KAM) theory
for the stability of invariant tori. 
The interlayer hopping is weak and short ranged and the angles are chosen in a large measure set. The result provides a justification, in the above regime,
to the effective continuum description of TBG in which large momenta interlayer interactions are neglected.
\end{abstract} 
\maketitle


\section{Introduction}  

The discovery that at certain angles Twisted Bilayer Graphene (TBG) develops superconductivity \cite{a}
has generated much interest in such materials both for technological and theoretical reasons \cite{b}. It was predicted, using continuum models obtained by keeping only the dominant harmonics in the lattice model
\cite{1}, \cite{2},\cite{3},  
that at such angles some strongly correlated behaviour should appear, but not of superconducting type. The mechanism behind the superconductivity remains elusive.

Taking the lattice into account breaks several symmetries of the continuum description 
\cite{1aa1}, \cite{1aa2} leading to effects like the possible shift of Fermi points.
More interestingly,
for generic angles, excluding a special set
\cite{1bb}, \cite{1aaa}, one has an incommensurate structure;
in such a case Bloch band theory does not apply  and one has
an emergent quasi-periodicity \cite{H1}-\cite{H5}, with some feature in common  with the one in fermions with quasi-periodic potentials
\cite{P0}-\cite{P2}. 
It is known that electronic quasi-periodic systems have remarkable properties. In 1d they produce a
metal-insulator transition \cite{Au},\cite{DS},\cite{FS}. The interplay with a many body interaction 
produces peculiar phases with anomalous gaps or many body localization  
\cite{Q0}--\cite{M11}. Quasi-periodicity has been studied also in Weyl semimetals 
\cite{P4}, \cite{W0},\cite{W1} or in the 2d Ising model \cite{Lu},
 \cite{Ca}, \cite{Ga}. 
It is therefore natural to expect that quasi-periodicity plays an important role in the interacting phases of TBG. 
Most theoretical analyis are however based on continuum effective description,
do not distinguish between commensurate and incommensurate angles, and are
based on the assumption that lattice effects
preserve the semimetallic phase \cite{1}, \cite{2}, \cite{3}.


We consider  
a lattice model for TBG
consisting of two graphene layers
one on top of the other and rotated by an angle $\th$. The momenta involved in the two-particle scattering process are of the form
$
 k_1-k_2+G+G'=0$ with $G=l_1b_1+l_2b_2\equiv lb$, $G'=m_1b'_1+m_2b_2'\equiv mb'$,
$l\equiv(l_1,l_2)\in\ZZZ^2$, $m\equiv(m_1,m_2)\in \ZZZ^2$, and $b_1,b_2$ are the vectors of the reciprocal lattice and $b'_i=R^T(\th) b_i$ the reciprocal lattice of the twisted layer
in which $R(\th)$ is the rotation matrix; the terms involving non zero $G,G'$ are also known as Umklapp interactions.
Note that, apart from special angles, $G'$ is not commensurate with $G$ and the effect of the mismatch of the lattices
is very similar to the effect of a quasi-periodic potential. This is quite clear  comparing for instance with the conservation law of 1d fermions
with Aubry-Andr\'e potential $\cos 2\pi \o x$, which is
$k_1-k_1+2l\pi+2\pi \o m=0$ with $\o$ irrational. 

It is expected that the relevant processes in TBG are the ones connecting the Dirac points as closely as possible, that is
the terms that minimize the quantity
$| G+G'+ p_{F,i}-p'_{F,j}|$ where $p_{F,i}$ $p'_{F,j}$ are the Dirac points
of the two layers. 
The approximation at the basis of the effective models \cite{1}, \cite{2},\cite{3}
consists in taking restricting the interaction to only the terms 
$G=G'=0$ or $G=b_1, G'=-b'_1$ or $G=b_2, G'=-b'_2$ and taking the continuum limit, based on the fact that larger
values of $G$ or $G'$ are exponentially depressed \cite{111}. 
However in the incommensurate case, Umklapp terms with
very large values of $G,G'$ make $| G+G'+ p_{F,i}-p'_{F,j}|$ arbitrarely small, producing almost relevant processes
which can destroy the semimetallic behaviour. 
In the 1d Aubry-Andr\'e model the processes that produce small values for
 $2 \e p_F+2l\pi+2\pi \o m$, $\e=0,\pm 1$ are indeed the ones producing the insulating behaviour at large coupling, while
at weak coupling the metallic regime persists. Similarly
the persistence or not of the semimetallic regime in TBG depends on the relevance or irrelevance of the terms involving large $G, G'$ that
almost connect the Dirac points. This fact cannot be understood only on the basis of perturbative arguments; it is indeed a non perturbative phenomenon which can be established only by the convergence or divergence of the whole series expansion. 
Despite the similarity of quasi-periodic potentials and incommensurate TBG, there are crucial differences like the higher dimensionality of TBG and the
fact that the frequencies are not independent parameters but are functions of a single parameter, the angle
between the layers, and this produces rather different small divisors.

The aim of this paper is to investigate when the quasi-metallic phase is stable against the large momentum processes
in the incommensurate case. The analysis is based on
Renormalization Group 
methods combined with number theoretical properties of irrationals,  similar to the ones used in Kolmogorov-Arnold-Moser (KAM) theory
for the stability of invariant tori. 
Due to the difficulty of getting information on the single particle spectrum, we analyze the
behavior of the Euclidean correlations, which provide information on the spectrum close to the Fermi points.
Such methods are robust enough to be extended to many-body systems, as it was done for the interacting Aubry-Andr\'e model \cite{M11}
or in Weyl semimetals \cite{W1}. Our main result is the proof of the stability
of the semimetallic phase in a large measure set of angles in the incommensurate case.

The paper is organized in the following way. In Section \ref{sec:model} the lattice model of TBG
is presented. In Section \ref{sec:feynman} a perturbative expansion for the correlations is derived.
In Section \ref{sec:feynman1} the emerging quasi-periodicity and the small divisor problem is described, together with the required (number theoretical) Diophantine conditions. Section \ref{sec:result} contains a statement of the main result and in Section \ref{sec:renormalized} the 
Renormalization Group derivation is presented. The Appendices detail the more technical aspects of the analysis.

\section{Incommensurate TBG}\label{sec:model}
We consider the lattice TBG model introduced in \cite{1},  \cite{2}.
We focus on this model for the sake of definiteness but our methods could be applied more generally.
We consider two graphene layers rotated with respect to one another by an angle $\theta$ around a point $\xi=(0,1/2)$ (that is, the point between an a and b atom, chosen so that the twisted model preserves the $C_2T$ symmetry in Appendix \ref{sec:symmetry}).
The Hamiltonian of the system will be written as
\begin{equation}
  H=H_1+H_2+V
\end{equation}
where $H_1$ and $H_2$ are hopping Hamiltonians within the layers 1 and 2 respectively and $V$ is an interlayer hopping term.
The first graphene layer is defined on the
lattice $\mathcal L_1:=\{n_1 A_1+n_2 A_2,\ n_1,n_2\in\mathbb Z\}$
with $A_1={1\over 2}(3,\sqrt{3})
  ,\quad
  A_2={1\over 2}(3,-\sqrt{3})$.
We introduce the nearest-neighbor vectors: $\d_1=(1,0)$, $\d_2={1\over 2}(-1, \sqrt{3})$, $\d_2={1\over 2}(-1, -\sqrt{3})$.
We will write the Hamiltonian in second quantized form: for $x\in \mathcal L_1$, we introduce {\it annihilation operators} $c_{1,x,a}$ and $c_{1,x,b}$ corresponding respectively to annihilating a fermion located at $x$ and $x+\d_1$.
The nearest neighbor hopping Hamiltonian is
\be H_1= -t\sum_{x\in \mathcal L_1}\sum_{i=0}^2 (c_{1,x,a}^\dagger c_{1,x+A_i,b}+ c_{1,x+A_i,b}^\dagger c_{1,x,a})\ee
where $A_0:=0$ (note that $\d_1-\d_2=A_2, \d_2-\d_3=A_3$).
We will do much of the computation in Fourier space, and we here introduce the Fourier transform $\hat c_{1,k,\alpha}$ of $c_{1,x,\alpha}^\pm$ in such a way that, for $\alpha\in\{a,b\}$,
\begin{equation}
  c_{1,x,\alpha}
  =\frac1{|\hat{\mathcal L}_1|}\int_{\hat{\mathcal L}_1} dk\ e^{-ik(x-\xi)}\hat c_{1,k,\alpha}
\end{equation}
with  $|\hat{\mathcal L}_1|=8 \pi^2/3 \sqrt{3}$, and
$
  \hat{\mathcal L}_1:=\mathbb R^2/(b_1\mathbb Z+b_2\mathbb Z)$ in which
\begin{equation}
  b_1= {\textstyle{2\pi\over 3}}(1,\sqrt{3})
 ,\quad
 b_2= {\textstyle{2\pi\over 3}}(1,-\sqrt{3})
 .
 \label{bi}
\end{equation}
In Fourier space,
\begin{equation}
  H_1= 
\frac t{|\hat{\mathcal L}_1|}\int_{\hat{\mathcal L}_1}dk\ \left(\Omega(k)\hat c_{1,k,a}^\dagger\hat c_{1,k,b}+\Omega^*(k)\hat c_{1,k,b}^\dagger\hat c_{1,k,a}\right)
\label{H1k}
\end{equation}
with $\Omega(k_x,k_y):=1+2e^{-i\frac32k_x}\cos({\textstyle\frac{\sqrt 3}2k_y})$. 


The second graphene layer is rotated by an angle $\theta$ around the point $\xi=(0,1/2)$, that is, it is defined on the lattice
\be \mathcal L_2=\xi+R(\theta)(\mathcal L_1-\xi),\quad R(\th)=\begin{pmatrix} c_\th &-s_\th \\s_\th & c_\th \end {pmatrix}\ee
(we use the shorthand throughout this paper that $c_\th\equiv\cos\th, s_\th\equiv\sin \th$).
The annihilation operators in the second layer are denoted by $c_{2,x,a}$ and $c_{2,x,b}$.
The hopping Hamiltonian of this second layer is
\begin{equation}
  H_2= -t\sum_{x\in \mathcal L_2}\sum_{i=0}^2 (c_{2,x,a}^\dagger c_{2,x+RA_i,b}+ c_{2,x+R A_i,b}^\dagger c_{2,x,a})
\end{equation}
where $R\equiv R(\theta)$.
We define the Fourier transform in the second layer: if
$b'_1:=R b_1$, $b'_2:=R b_2$
and
\begin{equation}
  c_{2,x,\alpha}
  =\frac1{|\hat{\mathcal L}_1|}\int_{\hat{\mathcal L}_2} dk\ e^{-ik(x-\xi)}\hat c_{2,k,\alpha}
\end{equation}
we find
\begin{equation}
\begin{array}{r@{\ }>\displaystyle l}
H_2=&
\frac t{|\hat{\mathcal L}_1|}\int_{\hat{\mathcal L}_2}d k\ 
\cdot\\&\cdot\left(\Omega(R^T k)\hat c_{2,k,a}^\dagger\hat c_{2,k,b}+\Omega^*(R^Tk)\hat c_{2,k,b}^\dagger\hat c_{2,k,a}\right)
.
\label{H2k}
\end{array}
\end{equation}

In the absence of interlayer coupling the two graphene layers are decoupled;
the single particle spectrum for layer 1 is $\pm |\Omega(k)|$
and the Fermi points
are given by the relation
$\O(p_{F,1}^\pm)=0$ with
\begin{equation}
  p_{F,1}^\pm={2\pi\over 3}(1,\pm  {\textstyle{1\over \sqrt{3}}})
  \label{pF}
\end{equation}
for momenta close to such points one has 
$|\Omega(k)| \sim {3\over 2} t  |k-p_{F,1}^\pm|$, that is the dispersion relation is 
almost linear (relativistic) up to quadratic corrections, forming approximate {\it Dirac cones}. In the same way the 
dispersion relation for layer 2 is $\pm |\Omega(R^T k)|$; 
the Fermi points are $\O(R^T p_{F,2}^\pm)=0$ with
$p_{F,2}^\pm=R(p_{F,1}^\pm)$
and $|\Omega(R^T k)| \sim {3\over 2} t  |k-p_{F,2}^\pm|$. We are interested in understanding how these four Dirac cones are modified in the presence of the interlayer hopping. 


We couple the 2 layers by an interlayer hopping Hamiltonian, which couples atoms of type a to atoms of type b:
\begin{equation}
\begin{array}{>\displaystyle l}
V=
\l \sum_{x_1\in \mathcal L_1} \sum_{x'_2\in\mathcal L_2}\sum_{\alpha\in\{a,b\}} \varsigma(x_1+d_\alpha-x'_2-Rd_{\alpha})
\cdot\\
\hfill\cdot(c^\dagger_{1,x_1,\alpha}c_{2,x'_2,\alpha}+c_{2,x_2',\alpha}^\dagger c_{1,x_1,\alpha})
\end{array}\end{equation}
$d_a=(0,0), d_b=\d_1$, and $\varsigma(x)=\varsigma(-x)$,
\be
\varsigma(x_1-x_2)=\int_{\mathbb R^2}\frac{dq}{4\pi^2}\  e^{i q(x_1-x_2)} \hat \varsigma(q)
,\quad
|\hat\varsigma(q)|\le e^{-\k |q|}
.
\label{interlayer}
\ee
We restrict the interlayer term to hoppings between atoms of type a to atoms of type a and type b to b for technical reasons.
This is so that the ``Inversion'' symmetry in Appendix \ref{sec:symmetry} is satisfied.
We could just as easily consider a model where the interlayer hopping occurs only between atoms of type a to atoms of type b.
We could relax this restriction and allow for all possible hoppings, but we would then need to add extra counterterms (see (\ref{counterterms})) to the model.
For the sake of simplicity, we avoid this and only consider these interlayer hoppings.


Note that, whereas the Fourier transform for $c$ is defined on $\hat{\mathcal L}_i$, the Fourier transform of $\varsigma$ is defined on all $\mathbb R^2$.
We write $V$ in Fourier space: we get, see App. \ref{app:fourierV}
\bea
&&V=
\frac \l{4\pi^2|\hat{\mathcal L}_1|}\sum_{\alpha}\left(\sum_{l\in\mathbb Z^2}\int_{\hat{\mathcal L}_1}d k \ 
\tau^{(1)}_{l,\alpha}(k+l b)
\hat c^\dagger_{1,k,\alpha}\hat c_{2,k+l b,\alpha}
\right.\nn\\
&&\left.+\sum_{m\in\mathbb Z^2}\int_{\hat{\mathcal L}_2}d k\
\tau_{m,\alpha}^{(2)}(k+m b') 
\hat c^\dagger_{2,k,\alpha}\hat c_{1,k+m b',\alpha}\right)
\nn\label{11}
\eea
where we use the notation $lb\equiv l_1b_1+l_2b_2$, $mb'\equiv m_1b_1'+m_2b_2'$, and
\begin{equation}
\tau^{(1)}_{l,\alpha}(k)
  :=e^{i \xi lb}e^{-ik(d_\alpha-Rd_{\alpha})}
  e^{-i\xi \sigma_{k,1}b'}\hat\varsigma^*(k)
  \label{tau1}
  \end{equation}
\begin{equation}
\tau^{(2)}_{m,\alpha}(k):=e^{i\xi mb'}e^{-ik(d_\alpha-Rd_{\alpha})}e^{-i\xi \sigma_{k,2}b}\hat\varsigma(k)
  \label{tau2}
\end{equation}
in which $\sigma_{k,i}\in \mathbb Z^2$ is the unique integer vector such that
$
  k-\sigma_{k,1}b'\in\hat{\mathcal L}_2
  ,\ 
  k-\sigma_{k,2}b\in\hat{\mathcal L}_1
  $. Note that the difference of the momenta of the two fermions is given by $l b+m b'$.










The position of the Dirac points are in general modified (renormalized)
by the interlayer hopping.
It is conventient to fix the values of the renormalized
Dirac points by properly choosing the bare ones. This can be achieved by replacing
$\O(k)$
with $\O(k)+\nu_{i,\o}$ close to each Dirac points, that is adding
a counterterm has the form
\begin{eqnarray}
&&M=
  \sum_{\omega\in\{+,-\}}\sum_{i=1,2}\int_{\hat{\mathcal L}_i} dk\ \chi_{\omega,i}(k)(  \nu_{i,\omega} \hat c_{i,k,a}^\dagger\hat c_{i,k,b}\nn\\
&&+ \nu^*_{i,\omega} \hat c_{i,k,b}^\dagger\hat c_{i,k,a})
\label{counterterms}
\end{eqnarray}
where $\chi_{\o,i}(k)$ is a smooth compactly supported function  that is
non vanishing for  $||k-p_{F,i}^\o||_i\le 1/\gamma$, for some $\gamma>1$, in which $||.||_i$ is the norm on the torus $\hat{\mathcal L}_i$. 


Our main result concerns the two-point Schwinger function, which we define as follows.
We first introduce a Euclidean time component: given an inverse temperature $\beta>0$, we define for $x_0\in[0,\beta)$,
\begin{equation}
  c_{j,x,\alpha}(x_0):=e^{-x_0\bar H}c_{j,x,\alpha}e^{x_0\bar H}
  .
\end{equation}
and $\bar H=H+M$. Combining the Euclidean time component with the spatial one, we define $\Lambda_i:=[0,\infty)\times \mathcal L_i$.
The corresponding Fourier-space operators are
\begin{equation}
  \hat
  c_{j,k,\alpha}(k_0)=
  \int_0^\beta dx_0 e^{-ix_0k_0}\sum_{x\in \mathcal L_j}e^{-i(x-\xi)k}c_{j,x,\alpha}(x_0)
\end{equation}
which is defined for $(k_0,k)\in\hat \Lambda_j:=\frac{2\pi}\beta \mathbb Z\times \hat{\mathcal L}_j$.

Now, given $j,j'\in\{1,2\}$, $\mathbf k=(k_0,k)\in \Lambda_j$, the two-point Schwinger function is defined as the $2\times2$ matrix $S_{j,j'}(\mathbf k)$ whose components are indexed by $\alpha,\alpha'\in\{a,b\}$:
\begin{equation}\label{xx}
  (\hat S_{j,j'}(\mathbf k))_{\alpha,\alpha'}:=
  \frac{\mathrm{Tr}(e^{-\beta \bar H}T(\hat c_{j,k,\alpha}(k_0),\hat c_{j',k,\alpha'}^\dagger(k_0)))}{\mathrm{Tr}(e^{-\beta\bar H})}
\end{equation}
where $T$ is the time ordering operator, which is bilinear, and is defined in real-space $
  T(c_{j,x,\alpha}(x_0),c_{j',y,\alpha'}^\dagger(y_0))=$
\begin{equation}
  \left\{\begin{array}{>\displaystyle ll}
    c_{j,x,\alpha}(x_0)c_{j',y,\alpha'}^\dagger(y_0)&\mathrm{if\ }x_0<y_0\\
    -c_{j',y,\alpha'}^\dagger(y_0)c_{j,x,\alpha}(x_0)&\mathrm{if\ }x_0\geqslant y_0
    .
  \end{array}\right.
\end{equation}


To compute the Schwinger functions, we will use the Grassmann integral formalism.
To do so, we add an imaginary time component to all position vectors: we define $\mathbf A_i:=(0,A_i)$.
Given $\mathbf x=(x_0,x)\in \Lambda_j\equiv [0,\beta)\times \mathcal L_j$, we introduce Grassmann variables
 $\psi^\pm_{\mathbf x,a,j}, \psi^\pm_{\mathbf x,b,j}$ and their Fourier transforms
\begin{equation}
  \psi_{j,\mathbf x,\alpha}^\pm
  =\frac1{|\hat\Lambda_j|}\int_{\hat\Lambda_j} d\mathbf k\ e^{\pm i\mathbf k(\mathbf x-\bm\xi)}\hat\psi_{j,\mathbf k,\alpha}^\pm
\end{equation}
with  $|\hat\Lambda_j|=8 \pi^2/3 \sqrt{3}$,
$\bm\xi=(0,\xi)$ and $\hat\Lambda_j=\frac{2\pi}\beta\mathbb Z\times \hat{\mathcal L}_j$. The Schwinger fiunctions are given by
\begin{equation}
  (S_{j,j'}(\xx,\yy))_{\alpha,\alpha'}:={\int P(d\psi)\ e^{-\b (V(\psi)+M(\psi))}  \psi^-_{j,\xx,\alpha} \psi^+_{j',\yy\,\alpha'}\over 
  \int P(d\psi)\ e^{-\b (V(\psi)+M(\psi))}}
\end{equation}
where $P(d\psi)=P(d\psi_1)P(d\psi_2)$ where $P(d\psi_1)$ is the Grassmann integration with propagator
\begin{equation}
  \hat g_1(\kk)=\begin{pmatrix} -i k_0 &\O(k) \\
  \O^*(k) & -i k_0\end{pmatrix}^{-1}
  .
\end{equation}
and $g_1(\xx,\yy)=\frac t{|\hat{\mathcal L}_1|}\int_{\hat\Lambda_1}d\mathbf k e^{i\kk(\xx-\yy)} \hat g_1(\kk)$
and 
\begin{equation}
 \hat g_1(\kk+p_F^\pm)\sim \begin{pmatrix} -i k_0  & -v_F(-ik_1\pm k_2) \\
  -v_F (i k_1\pm k_2) & -i k_0\end{pmatrix}^{-1}
  .
\end{equation}
$P(d\psi_2)$ has propagator
\begin{equation}
  g_2(\xx,\yy)={1\over |\hat{\mathcal L}_1|}\int_{\hat\Lambda_2}d\kk e^{i\kk(\xx-\yy)} \hat g_2(\kk)
\end{equation}
with $
  \hat g_2(\kk)=\begin{pmatrix} -i k_0  &\O(R^T k) \\
  \O^*(R^T k) & -i k_0\end{pmatrix}^{-1}
  \equiv
  \hat g_1(R^T \mathbf k)$
which is singular at $p_{F,2}^\pm:=R p_{F,1}^\pm$ and periodic in $k$ with period $b'_1, b'_2$.
The interaction and counterterm are rewritten formally in terms of Grassmann variables:
\bea\label{112}
V(\psi)=&&
\frac \l{4\pi^2|\hat\Lambda_1|}\sum_{\alpha}\\
&&\cdot 
\left(\sum_{l\in\mathbb Z^2}\int_{\hat\L_1}d \mathbf k \ 
\tau^{(1)}_{l,\alpha}(k+l b)
\hat\psi^+_{1,\mathbf k,\alpha}\hat\psi^-_{2,\mathbf k+l \mathbf b,\alpha}
\right.\nn\\
&&\left.+\sum_{m\in\mathbb Z^2}\int_{\hat\L_2}d\mathbf k\
\tau_{m,\alpha}^{(2)}(k+m b') 
\hat\psi^+_{2,\mathbf k,\alpha}\hat\psi^-_{1,\mathbf k+m \mathbf b',\alpha}\right)
\nn
\eea

\be
M(\psi)=\sum_{\o,\i}
\int   d\kk\ \chi_{\o,i}(k)
(\n_{i,\o}\psi^+_{i,\kk,\a}\psi^-_{i,\kk,\b}+\n^*_{i,\o}\psi^+_{i,\kk,\b}\psi^-_{i,\kk,\a})
\ee




\section{Perturbative expansion and small divisors}
\label{sec:feynman}


To compute the Schwinger functions $S(\mathbf k)$, we will use Feynman graph expansions. We give the rule setting $\n=0$ for definiteness.
The graphs for this model are chain graphs of the form depicted in Figure \ref{fig:feynman}

\begin{figure}
\hfil\includegraphics[width=8cm]{feynman.pdf}
\caption{\label{fig:feynman} Example of a Feynman diagram for $S_{1,1}(\mathbf k)$ with 4 vertices.}
\end{figure}


Each line $s$ is associated a layer label $i_s\in\{1,2\}$ and two valley indices $\alpha_s,\alpha'_s\in\{a,b\}$.
Each vertex has one entering line $s_1$ and an exiting line $s_2$, and we impose that these lines have different indices: $i_{s_1}=3-i_{s_2}$, and the same valley indices $\alpha_{s_1}=\alpha_{s_2}$.
To each vertex $r$ that has an entering line $s_1$ and exiting line $s_2$, we associate an index, which if $i_{s_1}=1$ we denote by $l_s\in \mathbb Z^2$, and if $i_{s_2}=2$ we denote by $m_s\in \mathbb Z^2$.
\begin{itemize}
  \item
  Each internal line $s$ with layer label $1$ coming from a vertex with index $m_s$ corresponds to the propagator $g_{1;\alpha_s,\alpha'_s}(\mathbf k+m_rb')$ (where $\mathbf k$ is the momentum at which $S$ is being evaluated).
  Each internal line $s$ with layer label $2$ coming from a vertex with index $l_s$ corresponds to the propagator $g_{2;\alpha_s,\alpha'_s}(\mathbf k+l_rb)$.
  \item
  Each external line $s$ corresponds to the propagator $g_{i_s:\alpha_s,\alpha'_s}(\mathbf k)$.
  \item
  Each vertex $r$ that has an entering line $s_1$ and exiting line $s_2$ with
  $i_{s_1}=1$ and $i_{s_2}=2$, if $s_1$ comes from a vertex with index $m_{r-1}$, the vertex $r$ corresponds to an interaction term $\lambda\tau_{l_r,\alpha_{s_2},\alpha'_{s_1}}^{(1)}(\mathbf k+m_{r-1}b'+l_rb)$.
  \item
  Each vertex $r$ that has an entering line $s_1$ and exiting line $s_2$ with
  $i_{s_1}=2$ and $i_{s_2}=1$, if $s_2$ comes from a vertex with index $l_{r-1}$, the vertex $r$ corresponds to an interaction term $\lambda\tau_{m_r,\alpha_{s_2},\alpha'_{s_1}}^{(2)}(\mathbf k+l_{r-1}b+m_rb')$.
\end{itemize}
The value of a graph is obtained by taking the product over the lines of the corresponding propagators, and multiplying them by the product over the vertices of the interaction terms.
The Schwinger function is then obtained by taking a sum over all possible graphs with the correct external labels.
A more explicit expression for the Schwinger function is given in Appendix \ref{app:explicit_feynman}.


The persistence or not of the semimetallic behaviour depends on the convergence of the expansion.
One of the more important and difficult reasons for which convergence could, in principle, not occur, is that small divisors could accumulate, as we will now explain.
Note that if $\kk_1$ and $\kk_2$ are 2 neighboring terms their difference is given by $k_1-k_2+m b'+l b$;
moreover the propagator are singular at the Dirac points $p_F^{\o, i}$. 
Consider therefore  $g_1(\kk) \t^{(1)}_l (k+l b) g_2(\kk+l b)$
and suppose that $k$ is in the first Brillouin zone and is close to 
$p_{F,1}^\o$, that is $k=p_{F,1}^\o+r_1$ with $r_1=O(\e)$ and $\e$ is a small parameter 
(so that the propagator is $O(\e^{-1})$);
suppose now that $k+l b+mb'$ is also in the first Brillouin zone (by periodicity we can always add $m b'$ to achieve this) 
and close to $p_{F,2}^{\o'}$, that is $k+l b+m b'=p_{F,2}^\o+r_2$ with $r_2=O(\e)$.
In principle this would produce an $O(\e^{-2} )$ contribution, that is an accumulation of small divisors which
could produce large contributions which can destroy convergence.
Note however that
\be
O(\e)=|r_1|+|r_2|\ge |r_1-r_2|=|l b+m b'-p_{F,2}^{\o'}-p_{F,1}^{\o}|\label{d1}
\ee
so that this accumulation of small divisors is possible only when the quantity in the r.h.s. is small.
The effect of such terms is rather delicate to be understoord.
Indeed 
similar terms in the case of random potential produce a localization phase in 1d,
while an extended phase in hgher dimension, while a quasi periodic disorder in 1d produces an extended phase.

\section{Diophantine condtions}
\label{sec:feynman1}
As we noted above, the accumulation of small divisors can only occur when
$|l b+m b'+p_{F,i}^{\o}-p_{F,j}^{\o'}|$ is small.
For generic values of $\theta$ (for which $s_\theta$ or $c_\theta$ is irrational), this happens when $l,m$ are large enough, and the basic issue for stability is if small divisors are balanced by the fact that the interlayer hopping is weak for large values of $l,m$.
We therefore need to quantify the relation between the size of $lb+mb'+p_{F,i}^\omega-p_{F,j}^{\omega'}$ and that of $l,m$, which we do using number theoretical considerations.
One such consideration is the so called 
{\it Diophantine} condition, which consists in restricting ourselves to values of $\theta$ for which, for any $\o, \o', i, j$
\be|p_{F,i}^\o-p_{F,j}^{\o'}+l b+m b'|\ge {C_0\over |y|^\t}\label{cond}
.
\ee
where $y$ is either $y=l$ or $y=m$ and $y\neq0$,
If this is satisfied, then when $|p_{F,i}^\o-p_{F,j}^{\o'}+l b+m b'|=O(\epsilon)$, then
\be
O(\e)\ge C_0 |l|^{-\t}.\ee
Therefore $|l|\ge \e^{-1/\t}$, and, due to the exponential decay of $\hat \varsigma$ in $\tau^{(1)}$ (see (\ref{interlayer})), so $|\t^{(1)}_l(k+l b)|\le e^{-\kappa\e^{-1/\t}}$ which compensates the small divisors $\e^{-2}$.  This simple argument says that small divisors due to adjacent propagators do not accumulate.
Of course this
argument is not sufficient to prove the convergence of the series, and then the persistence of cones;
it says only that two adjacent propagators cannot be simultaneously small but it says nothing about non adjacent ones, which is the generic case.
This requires a Renormalization Group analysis, as we will discuss below.

A basic preliminary question is whether there are $\th$'s such that \pref{cond} holds. Such condition is usually assumed 
in KAM theory or in the case of fermions with quasi periodic potentials; in such cases the frequencies are independent so that the fact that there is a large measure set of them verifying \pref{cond} follows from standard notions of number theory.
In the present case however the issue is much more subtle as 
all the frequencies depend on a single parameters $\th$. Nevertheless, we prove the following lemma.
\begin{lemma}\label{lemm:dioph}
For every interval $[\th_0,\th_1]\subset[0,2\pi)$, the set $\DD:=\{\th\in [\th_0,\th_1]$ satisfying \pref{cond}$\}$
has measure $1-O(C_0/(\th_1-\th_0)^2)$, so that choosing $C_0$ small enough with respect to $\th_1-\th_0$, $\mathcal D$
has large relative measure.
\end{lemma}

We will provide a full proof of this lemma in Appendix \ref{sec:dioph}.
For the moment, let us discuss the main idea of this proof.
In order to keep things simple, we will focus on the case $i=j$ and $\omega=\omega'$ here.
Let
$
M:=
lb+mb'$
and we wish to obtain a lower bound on $|M|$.
To do so, we use a simple inequality: $\forall x,y\in \mathbb R$,
\begin{equation}
  \sqrt{x^2+y^2}\ge
  \frac{\sqrt3}2x-\omega\frac12y
\end{equation}
and so, since
\begin{equation}
\begin{array}{rl}
M=
\frac{2\pi}3
(&l_1 +l_2 +m_1 \varphi_1  +m_2 \varphi_2
,\\&
\sqrt{3}
(l_1 -l_2 +m_1 \varphi_3  -m_2 \varphi_4)
)
\end{array}
\label{M}
\end{equation}
with $\varphi_1=c_\theta-s_\theta \sqrt{3}$, $\varphi_2=c_\theta+s_\theta \sqrt{3}$, $\varphi_3=c_\theta+s_\theta/ \sqrt{3}$, $\varphi_4=c_\theta-s_\theta/ \sqrt{3}$,
we have
\begin{equation}
  |M|\ge
  \frac{2\pi}{\sqrt3}(l_\omega+m\cdot f_\omega)
  \label{Mineq}
\end{equation}
with $l_+\equiv l_1$ and $l_-\equiv l_2$, and
\begin{equation}
  f_\omega(\theta):=({\textstyle \frac{\varphi_1-\omega\varphi_3}2},\ {\textstyle \frac{\varphi_2+\omega\varphi_4}2})
  .
\end{equation}
Thus, we wish to impose a condition on $\theta$ such that
$g_{l_\omega,m}(\theta):=| l_\omega
+m \cdot f_\omega(\theta)|\geqslant C_1|m|^{-\tau}$.
The measure of the complement of the 
set where this is true is bounded by
\be
\sum_{m,l_\omega}^* \int_{-C_1|m|^{-\tau}}^{C_1|m|^{-\tau}} \frac1{g_{l_\omega,m}'}dg_{l_\omega,m}\ee
where $g'_{l_\omega,m}$ is the derivative of $g_{l_\omega,m}$, and $\sum^*_{k,l}$ has the constraint that $\exists \theta\in[\theta_0,\theta_1]$ such that $g_{l_\omega,m}(\theta)\in[-C_1|k|^{-\tau},C_1|k|^{-\tau}]$. Therefore we get the bound, taking into account the sum over $l_\omega$,
\be
\sum^*_{m} 2C_1 {|m|^{1-\tau}\over \min_{\theta} |m\cdot f_\omega'(\theta)|}\label{kxx}
\ee
Now, in order for this bound to be useful, we need a good bound on $m\cdot f_\omega'$.
Now $m\cdot f_\omega'(\theta)$ can be small, but only if $m$ is large enough.
This suggests we should control it using another Diophantine condition, but we would run into an infinite problem: we would need $m f_\omega''$ to be bounded below, for which we would impose an extra Diophantine condition, which would itself require a Diophantine condition on the third derivatives, et c\ae tera.
We can avoid this problem as follows.
If
$f_\omega'(0)=|f_\omega'(0)|(\cos\b,\sin\b)$ is non vanishing and $\th$ is small then
$(m\cdot f_\omega'/|f_\omega'|)\sim |m|\cos(\th_m)$
where $\th_m$ is the angle between $\b$ and $m$.
We can distinguish in the sum (\ref{kxx}) the sum over $m$
such that $|\th_m|, |\th_m-\pi| < \pi/4$ and the complementary set
 $|\th_m-\pi/2|, |\th_m-3\pi/2| < \pi/4$.
In the first sum, $(m\cdot f_\omega'/|f_\omega'|)$ will be greater than $|m|$
up to some constant; we impose a Diophantine condition only for the second term:
for $|\th_m-\pi/2|, |\th_m-3\pi/2| < \pi/4$ we assume that $|l_\omega+m \cdot f_\omega'|\le C_1 |m|^{-\t}$.
Again we will end-up with a condition like (\ref{kxx})
involving 
$m\cdot (f_\omega'/|f_\omega'|)'$ which in principle could be arbitrarily small.
However $(f'/|f'|)'$
is orthogonal to $f_\omega'/|f_\omega'|$ hence 
$m\cdot (f_\omega'/|f_\omega'|)'\sim |m|\cos(\th_m+\pi/2)$ 
which, in this region, is greater than $|m|$.

Thus, we construct a large-measure set of $\theta$'s that satisfy $|l_\omega+m\cdot f_\omega(\theta)|\geqslant C_1|m|^{-\tau}$.
A detailed proof can be found in Appendix \ref{sec:dioph}.


\section{Stability of the semimetallic phase} \label{sec:result}

We are now ready to state our main result.
\vskip.2cm
{\bf Theorem}
{\it 
For any interval $[\theta_0,\theta_1]\subset [0,2\pi)$, any $C_0>0$,
there exists a subset of $[\theta_0,\theta_1]$ whose measure is at least $1-O(C_0/(\theta_0-\theta_1)^2)$ such that \pref{cond}
holds,
and an $\e_0$ (depending on $C_0,\th_0,\th_1$), such that, for any $|\l|\le \e_0$, for a suitable choice of $\n_{i,\o}$ 
$
  (\hat S_{j,j}(\mathbf k+\mathbf p_{F,j}^\o ))=$
\be
\begin{pmatrix}\label{43}
 -i Z_{j,\o} k_0  &  (i v_{j,\o} k_1- w_{j,\o} \o k_2 )\\
(-i v_{j,\o}^* k_1- w_{j,\o}^* \o k_2) &   -i Z_{j,\o} k_0
\end{pmatrix}^{-1}(1+O(|k|^{\alpha}))
\ee
with $0\le \alpha\le 1$, 
$Z=1+O(\l)$ real and 
$v_{i,\omega}=3t/2+O(\l)$, $w_{i,\omega}=3t/2 +O(\l)$, 
$\n_{j,\o}=O(\l)$. 
}

\vskip.2cm
This result ensures that, even taking into account the 
Umklapp
processes involving the exchange of very high momenta due to the emerging quasi-periodicity, the Weyl semimetallic phase persists
for small interalyer coupling and a large measure set of angles.

The interlayer coupling modifies
the position of the Dirac points; 
we have properly chosen the bare Dirac points $p_{F,i}^\pm(\l)$ in absence of interlayer coupling given by $\O(p_{F,i}^\pm(\l))+\n_{i,\pm}=0$)
so that their renormalized physical value is $p_{F,i}^\pm$ given by (10).
This is essentially equivalent to say that the position of the Dirac points genericaly moves in a way depending on the angle, the layer and the coupling.

The velocities
$w_{j,\o}, v_{j,\o}$  and the wave function normalization $Z_{j,\o}$,
are renormalizated in a way generically dependent on the layer and the angle.
Note that a priori several other relevant terms could be present, but they are excluded by symmetry. The singularity
of the Schiwnger function is given by
$Z^2 k_0^2+R(k)$ with $R(k)\sim (|v|^2 k_1^2+|w|^2 k_2^2)
$; the singularity of the 2-point function is therefore the same as in absence of interlayer at weak coupling ensuring the stability of the semimetallic phase.

The result holds for irrational twisting angles verifying \pref{cond}. The relative measure of this set can be made arbitrarely close to $1$
by decreasing $C_0$. In the remaining sections
we prove the above result by a Renormalization Group analysis, leading
to a convergent expansion.


\section{The renormalized expansion}\label{sec:renormalized}

\subsection{Multiscale decomposition}
\label{sec:multiscale}


We introduce smooth cut-off functions: for $i=1,2$, $\o=\pm$, $h\in\{-\infty,\cdots,0\}$, let $\chi_{h,i,\o}(\kk)$ be a smooth function that vanishes outside the region $||\kk-\pp_{F,i}^\omega|| \le \g^{h-1}$ and that is equal to 1 for $||k-\pp_{F,i}^\omega||\ge \g^{h-2}$.
The constant $\g>1$ will be chosen below to be large enough. Note that, in this way, the supports of $\chi_{0,i,+}$ and $\chi_{0,i,-}$ do not overlap.
We define $\hat g_{i,\o}^{(\le 0)}(\kk)=\chi_{0,i,\o}(\kk)\hat g_i(\kk)$
and 
\be \hat g_i(\kk)= g_i^{(1)}(\kk)+\sum_{\o=\pm} \hat g^{(\le 0)}_{i,\o} (\kk)
\ee
with $\hat g^{(1)}(\kk)=(1-\sum_\o \chi_{0,i,\o}(k))\hat g_i(\kk)$; this induces the Grassmann variable decomposition $\hat\psi_{i,\kk,\alpha}=\hat\psi_{i,\kk,\alpha}^{(1)}+\sum_{\o=\pm} \hat\psi^{(\le 0)}_{i,\kk,\alpha,\o}$
with propagators given by  $\hat g^{(1)}_i$ and $\hat g^{(\le 0)}_{i,\o}$
respectively. Note that $\hat\psi^{(1)}$ correspond to fermions with momenta far from the Fermi points, while  $\hat\psi^{(\le 0)}$ with momenta around $\pm \pp_{F,i}$.

 
We further decompose 
\be
\hat g_{i,\o}^{(\le 0)} (\kk)=\sum_{h=-\io}^0 
\hat g^{(h)}_{i,\o}( \kk)
\ee
where $\hat g_{i,\o}^{(h)}(\kk):=f_{h,i,\o}(\kk)\hat g_{i,\o}^{(\le 0)}$ in which $f_{h,i,\o}:=\chi_{h,i,\o}-\chi_{h-1,i,\o}$ is a smooth cutoff function supported in $ \g^{h-3} \le |\kk-\pp^\o_{F,i}|\le  \g^{h-1}$ such that $\sum_{h=-\io}^0f_{h,i,\o}=\chi_{0,i,\o}$. 
The integration is done recursively in the following way: assume that we have integrated the fields $\psi^{(1)},..,\psi^{(h-1)}$ obtaining
\be
e^{W}=\int\bar P(d\psi^{(\le h)}) e^{V^{(h)}(\psi,\phi)}  
\ee
where $\bar P(d\psi^{(\le h)})$ is Gaussian integration with propagator $\bar g_{i,\omega}^{(\le h)}$ which will be defined inductively in (\ref{prop_ind}),
and
\bea
&&V^{(h)}(\psi,0)=
\\
&&\sum_{i,\o,\o',l,\alpha,\alpha'} \int_{\hat\L_i} d\kk W^{(h,\omega,\omega')}_{i,2,l,\alpha,\alpha'}(\kk) 
\psi_{i,\kk,\alpha,\o}^+\psi_{2,\kk+lb,\alpha',\o'}^-+\nn\\
&&\sum_{i,\o,\o',m,\alpha,\alpha'} \int_{\hat\L_i}  d\kk W^{(h,\omega,\omega')}_{i,1,m,\alpha,\alpha'}(\kk) 
\psi_{i,\kk,\alpha,\o}^+\psi_{1,\kk+mb',\alpha',\o'}^-
.
\label{eff}
\eea

According to the RG procedure, we renormalize the relevant and marginal terms; we will see below that the term
with $l$ or $m$ non zero are actually irrelevant, due to improvements in the estimates due to the Diophantine condition. 
We therefore define a localization operation in the following way
\bea
&&\LL W^{(h,\omega,\omega')}_{i,j,l}(\kk)=\d_ {\o,\o'}\d_{i,j} \d_ {l,0}[W^{(h,\omega,\omega)}_{i,i,0}(0,p_{F,i}^\o)+\nn\\
&&k_0 \partial_0 W^{(h,\omega,\omega)}_{i,i,0}(0,p_{F,i}^\o)+
(k-p_{F,i}^\o)\partial  W^{(h,\omega,\omega)}_{i,i,0}(0,p_{F,i}^\o)
.
\label{loc}
\eea
The terms for which $\LL=0$ are called {\it non resonant} terms and the ones for which $\LL\not=0$ {\it resonant} terms. The terms containing derivatives
are marginal ones and produce wave function or velocities renormalizations, while the terms without derivatives are the relevant terms. 
The action of $\mathcal L$ on the effective potential $V^{(h)}$ is
\begin{equation}
  \LL V^{(h)}=\LL_1 V^{(h)}+\LL_2 V^{(h)}
\end{equation}
with
\begin{equation}
\LL_1 V^{(h)}:=\sum_{i,\omega,\a,\a'}\int_{\hat \Lambda_i} d\kk \g^h\n_{h,\omega,\a,\a',i}
\psi^+_{\kk,\o,i,\a}\psi^-_{\kk,\o,i,\a'}
\end{equation}
and
\begin{equation}
\LL_2 V^{(h)}:=\sum_{i,j,\omega,\a,\a'}\int_{\hat \Lambda_i} d\kk
z_{h,\o,\a,\a',i,j} (\mathbf k-\mathbf p_{F,i}^\omega)_j\psi^+_{\kk,\o,i,\a} \psi^-_{\kk,\o,i,\a'}
\end{equation}
with
\begin{equation}\label{ai}
  \nu_{h,\omega,\alpha,\alpha',i}:=\gamma^{-h} W_{i,i,0,\alpha,\alpha'}^{h,\omega,\omega}(0,p_{F,i}^\omega)
\end{equation}
\begin{equation}
  z_{h,\omega,\alpha,\alpha',i,j}:=-\partial_{\mathbf k_j}W_{i,i,0,\alpha,\alpha'}^{h,\omega,\omega}(0,p_{F,i}^\omega)
  .
\end{equation}

The form of the resonant terms is severely constrained by symmetries: as is proved in Appendix \ref{sec:symm},
\bea
&&\n_{h,\o,a,a,i}=\n_{h,\o,b,b,i}=0\quad \n_{h,\o,a,b,i}=\n^*_{h,\o,b,a,i}\nn\\
&&z_{h,\o,b,a,i,1}=z_{h,\o,a,b,i,1}^*\quad
z_{h,\o,b,a,i,2}=z_{h,\o,a,b,i,2}^*
\\
&&z_{h,\o,a,a,i,1}=z_{h,\o,b,b,i,1}=
z_{h,\o,a,a,i,2}=z_{h,\o,b,b,i,2}=0
\nn
\\
&&z_{h,\o,b,a,i,0}=z_{h,\o,a,b,i,0}=0
\quad
z_{h,\o,a,a,i,0}=z_{h,\o,b,b,i,0}\in i \mathbb R
\nn
\eea

The contributions from $\mathcal L_2 V$ are marginal, and are absorbed into the propagator at every step of the integration:
\begin{equation}
  \begin{array}{>\displaystyle l}
  \bar g_{i,\omega}^{(\le h)}(\mathbf k)
  :=
  \chi_{h,i,\omega}(\mathbf k)
  \cdot\\\hfill\cdot
  \left((\bar g_{i,\omega}^{(\le h+1)}(\mathbf k))^{-1}
  -\sum_j z_{h,\omega,\cdot,\cdot,i,j}(\mathbf k-\mathbf p_{F,i}^\omega)_j\right)^{-1}
  \end{array}
  \label{prop_ind}
\end{equation}
Thus,
\begin{equation}
  \begin{array}{>\displaystyle l}
\bar g_{i,\o}^{(\le h)}(\kk+\pp_{F,i}^\o)=
\chi_{h,i,\o}(\kk)(1+O(k))
\cdot\\\hfill\cdot
\begin{pmatrix}
 -i Z_{i,\o,h} k_0  &  (i v_{i,\o,h} k_1- w_{i,\o,h} \o k_2)\\
(-i v^*_{i,\o,h} k_1- w^*_{i,\o,h} \o k_2) &   -i Z_{i,\o} k_0
\end{pmatrix}
^{-1}\label{prop}
\end{array}
\end{equation}
with
\bea
  &&Z_{i,\omega,h}=Z_{i,\omega,h+1}-iz_{h,\o,a,a,i,0}
  \nn\\
  &&v_{i,\omega,h}= v_{i,\omega,h+1}+i z_{h,\o,a,b,i,1}
  \nn\\
  &&w_{i,\omega,h}= w_{i,\omega,h+1}+ \omega z_{h,\o,a,b,i,2}
  .
\eea

After absorbing $\mathcal L_2V^{(h)}$ into the propagator, we are left with integrating $\LL_1 V^{(h)}$ and
\begin{equation}
  \RR V^{(h)}:=(1-\mathcal L)V^{(h)}
\end{equation}
so
\be
e^{W}=\int \bar P(d\psi^{(\le h)}) e^{\LL_1 V^{(h)}(\psi)+\RR V^{(h)}(\psi)}  
.
\label{renormalized_expansion}
\ee

\subsection{Feynman rules for the renormalized expansion} \label{sec:renormfeyn}

The renormalized expansion described above has a graphical representation that is similar to the Feynman diagram expansion from Section \ref{sec:feynman}.
There are two main differences: first, there are two different types of vertices: ``\emph{$\tau$-vertices}'', coming from $\mathcal R V^{(h)}$ in (\ref{renormalized_expansion}), and ``\emph{$\nu$-vertices}'', coming from $\mathcal L_1 V^{(h)}$.
Second, every line has a {\it scale label} $h$, corresponding to a propagator on scale $h$:
\begin{equation}
  \bar g_{i,\omega}^{(h)}(\mathbf k):=f_{h,i,\omega}(\mathbf k)\bar g^{(\le h)}_{i,\omega}(\mathbf k)
  .
\end{equation}

The scale labels induce an important structure: given a diagram, we group vertices together into nested \emph{clusters}, which are connected subgraphs in which the scales of the lines leaving the cluster are all smaller than the scales of the lines inside the cluster, see Figure \ref{fig:clusters}.
A cluster that is such that $\mathcal L$ applied to the cluster yields $0$ is called {\it non-resonant}, otherwise it is called {\it resonant}.
In other words, the action of $\mathcal R=1-\mathcal L$ is trivial on non-resonant clusters, and non-trivial on resonant ones.

Some clusters are single vertices (either $\nu$ or $\tau$) and are called \emph{trivial clusters}.
The clusters that contain internal lines are called \emph{non-trivial clusters}.
As per the construction above, if $h^{ext}_T$ is the largest of the scales of the external lines of a non-trivial cluster $T$, all its internal lines have a scale $h>h^{ext}_T$; $h_T$ is the largest scale of the propagators internal to the cluster $T$.
A non-trivial cluster $T$ contains sub-clusters $\tilde{T}\subset T$.
We call a cluster $\tilde T\subset T$ a \emph{maximal cluster}
if there is no other cluster $\bar T$ such that  $\tilde T\subset \bar T\subset T$.

For each cluster, there are two external lines that connect the cluster to other ones.
In a non-resonant cluster $T$ with external lines of type $i_1=i_2=1$ and momenta
$k_1, k_2$ with $k_1$ in the first Brillouin zone
and $k_2=k_1+\hat m_T b+l b$ where $l b$ is chosen so that $k_2$ is in the first Brillouin zone,
if $A_0,..,A_N$
are the momenta associated to the $\t$ vertices contained in $T$ (the $\nu$ vertices do not change momentum)
one has $A_0=k_1+ l_0 b$, $A_1=k_1+l_0 b+m_1 b'$, $A_2=k_1+m_1 b'+l_2 b$, $A_3=k_1+m_3 b'+l_2 b$,..., $A_{N}=k_1+m_{N} b'+l_{N-1} b$ and 
$m_{N}=\hat m_T$
with $N$ odd.
In the same way if the non-resonant cluster $T$ has one external line of type $i_1=1$ and momentum $k_1$, and one of type $i_2=2$ and momentum
$k_2$; assume that $k_1$ is in the first Brillouin zone
and  $k_2=k_1+\hat l_T b+m b'$,
where $m b'$ is chosen such that $k_2$ is in the first Brillouin zone.
Now with $N$ even 
the momenta associated to the $\t$ vertices in $T$ are
$A_0=k_1+ l_0 b$, $A_1=k_1+l_0 b+m_1 b'$, $A_2=k_1+m_1 b'+l_2 b$, $A_3=k_1+m_3 b'+l_2 b$,...$A_{N-1}=k+m_{N-2} b'+l_{N-1}b$,
$l_{N-1}=\hat l_T$.
 

The value associated to a graph $\G$ is denoted by
$W_\G(\kk)$ and is given by the product of the propagators and $\n,\t$ factors associated to the vertices, with the $\RR$ operation acting on each 
non-resonant cluster. The effect of the $\RR$ operation on the non-resonant clusters can be written as
\be
\RR W^h(\kk+p_{F,i}^\o)=k^2\int_0^1 \partial^2 W(t \kk)\ee 


\begin{figure}
\hfil\includegraphics[width=8cm]{cluster.pdf}

\caption{\label{fig:clusters} An example of graph of order $\l^7$ with the associated clusters, denoted by thick rectangles. In this example, $h<h_1<h_2<h_3$.}
\end{figure}



\subsection{Convergence of the series and persistence of the cones}


In order to bound the contribution of a Feynman graph we note that $|\bar g^{(h)}(k)|\le C \g^{-h}$ (by (\ref{prop})); therefore the product of propagators in the Feynman graph is bounded by
\be
\prod_{T\mathrm{\ n.t.}} \g^{-h_T (M_T+R_T-1)}\ee
where $M_T$ is the number of maximal non-resonant clusters contained in $T$, $R_T$ is the number of maximal resonant clusters contained in $T$, and $\prod_{T \mathrm{\ n.t.}}$ is a product over non-trivial clusters.
Note that the trivial clusters do not contribute, as they contain no internal lines.
The effect of the $\RR:= \mathds 1-\mathcal L$ (recall (\ref{loc})) operation on the non-trivial resonant clusters 
produces an extra $\g^{2 (h_{T}^{\text{ext}}-h_T)}$ (the gain $h_T^{\mathrm{ext}}$ comes from the extra $(k-p_{F,i}^\omega)^2$ terms (which are on the scale of the external lines) and the $h_T$ loss from the extra second derivative (which are on the scale of the internal lines)).
Therefore, we get an extra factor:
\be
\prod_{T\ \mathrm{res\ n.t.}} \gamma^{2(h_{T}^{\text{ext}}-h_{T})}
\ee
where $\prod_{T\ \mathrm{res\ n.t.}}$ is the product over the non-trivial resonant clusters.
In addition, each maximal resonant cluster corresponds to a weight $\gamma^h\nu_h$, so, if $|\nu_h| \leqslant C |\lambda|$, we get an extra factor:
\begin{equation}
  \prod_{T \mathrm{\ n.t.}}\gamma ^{h_T M_T^\nu} C |\lambda|
\end{equation}
where $M_T^\nu$ is the number of maximal resonant trivial clusters in $T$.
Thus, we get the following bound for the sum over all the labels of a renormalized graph with $q$ vertices:
\bea
&&\sum^*_{\underline h,\underline l, \underline m} [L(\underline l, \underline m)] 
|\l|^q C^q
\left( \prod_{T\, \text{n.t.}} \g^{-h_T (M_T+R_T-1)}
\right) \nn\\
&&\Bigg(\prod_{T\ \mathrm{res\ n.t.}} \gamma^{2(h_{T}^{\text{ext}}-h_{T})}\Bigg)
\prod_{T  \, \text{n.t.}} \gamma^{h_{ T} M^\n_{T}}
\label{lap} 
\eea
where $L(\underline l, \underline m)$ is the norm of the product of the $\t$ functions:
\begin{equation}
  L(\underline l,\underline m):=\prod_i| \t(A_i)|
\end{equation}
where $A_i$ are the momenta, which depend on the structure of the graph, and $\sum^*_{\underline h,\underline l, \underline m}$ is the sum over the scales and momenta associated to the vertices, as explained in Section \ref{sec:renormfeyn}; 
these sums are not independent, as they are constrained by the compact support properties of the propagators and of the Diophantine condition.
Now
        defining $\mathds 1_{\Gamma\,\mathrm{res}}$ is equal to 1 if the maximal cluster $T_m$ is resonant and 0 otherwise, we have
\bea
&&\prod_{T\, \text{n.t.}}  \g^{-h_T R_T}  \prod_{T\ \mathrm{res\ n.t.}, T\not=T_m} 
\gamma^{h_{T}^{\text{ext}}} \prod_{T\, \text{n.t.}}
\gamma^{h_{ T} M^\n_{T}}\le 1\nn\\\
&&\prod_{T\, \text{n.t.}}  \g^{h_T}
\prod_{T\ \mathrm{res\ n.t.}, T\not= T_m} 
\gamma^{-h_{T}}
\le \gamma^{h_{T_m} \mathds 1_{\Gamma\,\mathrm{res}}}
\eea 
and then using that 
if $T_m$ is resonant and $\LL$ is applied then $h_{T_m}=h+1$ we get
\be
\prod_{T\, \text{n.t.}}  \g^{-h_T (R_T-1)} \gamma^{h_{ T} M^\n_{T}}\prod_{T\ \mathrm{res\ n.t.}} 
\gamma^{(h_{T}^{\text{ext}}-h_{T})}
\le \gamma^{(h+1)\mathds 1_{\Gamma\,\mathrm{res}}}.\ee 

Therefore, \pref{lap} is bounded by 
\bea
&&\sum^*_{\underline h,\underline l, \underline m} [L(\underline l, \underline m)] 
|\l|^q C^q
\gamma^{h \mathds 1_{\Gamma\,\mathrm{res}}}
\nn\\&&
\left( \prod_{T\, \text{n.t.}} \g^{-h_T M_T}
\right)
\Bigg(\prod_{T\ \mathrm{res\ n.t.}} \gamma^{(h_{T}^{\text{ext}}-h_{T})}\Bigg)
.
\label{lap1} 
\eea
As the graphs are chains, there is no problematic combinatorial factor; the main issue
is the sum over $\underline h$; if we neglect the constraint in 
$\sum^*_{\underline h,\underline l, \underline m}$ then we will get a factor  $
\prod_{T\, \text{n.t.}} 
 (\sum_{h_T=-\io}^0 \g^{-h_T})$ which is divergent.

However, we have still not taken advantage of the Diophantine condition. In order to do that we note that 
$
|\t^{(i)}(k)|\le C e^{-\kappa |k|}
$
from the exponential decay or $\hat \varsigma(k)$, see (\ref{interlayer}), (\ref{tau1})-(\ref{tau2}); we can write 
\be
e^{-\kappa/2 |k|}=\prod_{h=-\io}^0 e^{-\kappa 2^h |k|}
\ee Consider the product of $\t$ factors:
\be
L(\underline l,\underline m)\equiv\prod_i| \t(A_i)|\le \prod_i e^{-\kappa |A_i|/2}
\prod_{h=-\infty}^1 e^{-\kappa 2^{h} |A_i|}\ee
which we estimate as
\be
L(\underline l,\underline m)\le \prod_i e^{-\kappa |A_i|/2}
\prod_{T\ \mathrm{nonres}\ni i} e^{-\kappa 2^{h_T} |A_i|}\ee
where in the last product $i$ is the label of the points $\t$ contained in $T$.
We then exchange the products:
\be
L(\underline l,\underline m)\le \left(\prod_i e^{-\kappa |A_i|/2}\right)
\prod_{T\ \mathrm{nonres}} \prod_{i\in T}e^{-\kappa 2^{h_T} |A_i|}\ee
Let us consider a non resonant cluster with external lines of type $1$;
we have
\be
|A_0|+|A_1|+\ldots\ge  |A_0-A_1+A_2-A_3\ldots|= |\hat m_T b'|\ee
(see Section \ref{sec:renormfeyn}) so that
\be
\prod_{i\in T}e^{-\kappa 2^{h_T} |A_i|}
\le e^{-\kappa 2^{h_T} |\hat m_T b'|}
\ee
A similar analysis holds for $i_1=i_2=2$ with $\hat l\not=0$.
In the same way if the cluster $T$ has $i_1=1$ and $i_2=2$ and 
$k_1, k_2$ are the momenta of the external lines
using that $|A_0|+|A_1|+\ldots$
\be\ge |A_0-A_1+A_2-A_3\ldots|\ge |k_1+\hat l b|\ge |\hat lb|-\frac{4\pi}3\ee
(where we used that $|k_1|\le\frac{4\pi}3$) as $k_1$ is in the first Brillouin zone
\be
\prod_{i\in T}e^{-\kappa 2^{h_T} |A_i|}\le
e^{-\kappa 2^{h_T} (|\hat l_T b|-\frac{4\pi}3)}
.
\ee
In addition, the Diophantine condition imposes a constraint between the momenta
of the external lines of a cluster $T$ and the $l,m$ labels associated to its internal vertices.
Consider a non resonant cluster with the following indices:
\begin{enumerate}
\item If $i_1=i_2=1$ and 
$k_1, k_2$ are the momenta of the external lines; assume that $k_1$
is in the first Brillouin zone
and $k_1=:\bar k_1+p_{F,1}^{\o_1}$. Moreover $k_2:=k_1+\hat m_T b'+l b=:\bar k_2+p_{F,1}^{\o_2}$,
with $l$ chosen such that $k_2$ is in the first Brillouin zone; then, if $\hat m_T\not=0$, by (\ref{cond}),
\bea
&&2 \g^{h^{ext}_T}\ge |\bar k_1|+|\bar k_2|\ge |\bar  k_1-\bar k_2|=\nn\\
&&|
p_{F,1}^{\o_1}-p_{F,1}^{\o_2}+\hat m_T b'+l b|\ge {C_0\over |\hat m_T|^\t}\eea
so that, if $\hat m_T\not=0$
\be
|\hat m_T|\ge ({\textstyle\frac12}C_0 \g^{-h^{\mathrm{ext}}_T})^{\frac1\tau}\label{cond1}
\ee
and so
\begin{equation}
  |\hat m_Tb'|\ge c_1\g^{-h^{\mathrm{ext}}_T/\tau}
\end{equation}
for some constant $c_1>0$.
On the other hand if $\hat m_T=0$ but $\o_1\not=\o_2$ then $l=0$ so $\g^{h^{\mathrm{ext}}_T}\ge\frac12|p_{F,1}^{\omega_1}-p_{F,1}^{\omega_2}|=\frac{2\pi}{3\sqrt3}$ (recalling (\ref{pF})).
Thus, this eventuality does not occur provided $\gamma$ is large enough.

\item
In the case $i_1=1$ and $i_2=2$ and 
$k_1, k_2$ are the momenta of the external lines; assume that $k_1$  is in the first Brillouin zone
and $k_1=:\bar k_1+p_{F,1}^{\o_1}$. Moreover $k_2:=k_1+\bar l b+m b'=:\bar k_2+p_{F,2}^{\o_2}$,
with $m$ chosen in such a way that $k_2$ is in the first Brillouin zone; then, if $\bar l\not=0$, by (\ref{cond}),
\bea
&&2 \g^{h^{ext}_T}\ge |\bar k_1|+|\bar k_2|\ge |\bar  k_1-\bar k_2|=\\
&&|p_{1,F}^{\o_1}-p_{2,F}^{\o_2}+\hat l_T b+m b'|\ge {C_0\over |\hat l_T|^\t}\nn\eea
so that
\be|\hat l_T|\ge ({\textstyle\frac12}C_0 \g^{-h^{\mathrm{ext}}_T})^{\frac1\tau}\label{cond2}
\ee
and so
\begin{equation}
  |\hat l_Tb|\ge c_1\g^{-h^{\mathrm{ext}}_T/\tau}
\end{equation}
 If $\hat l_T=0$ then $2 \g^{h^{\mathrm{ext}}_T}\ge O(\th)$ for $\o_1=\o_2$ and $2 \g^{h^{\mathrm{ext}}_T}\ge O(1)$ for $\o_1=-\o_2$.
Thus, provided $\gamma\gg \theta^{-1}$, these eventualities do not present themselves provided $\gamma$ is large enough.

\item
A similar analysis holds for $i_1=2, i_1=1$, and $i_1=i_2=2$.
\end{enumerate}

Thus,
\begin{equation}
  \prod_{i\in T}e^{-\kappa 2^{h_T}|A_i|}\le e^{-\kappa 2^{h_T}(c_1 \gamma^{-h_T^{\mathrm{ext}}/\tau}-\frac{4\pi}3)}
\end{equation}
which, provided $\gamma$ is large enough, yields
\begin{equation}
  \prod_{i\in T}e^{-\kappa 2^{h_T}|A_i|}\le e^{-c_2\gamma^{-h_T^{\mathrm{ext}}/\tau}}
\end{equation}
for some constant $c_2$.
Therefore,
\be
L(\underline l,\underline m)\le e^{-c_2 \gamma^{-h/\tau}\mathds 1_{\Gamma\,\mathrm{nonres}}}
\prod_i e^{-\kappa |A_i|/2}
\prod\limits_{T\ \mathrm{n.t.}}
e^{-c_2 {M}_{T} \gamma^{-h_{T}/\tau}}\ee
where $\mathds 1_{\G\,\mathrm{nonres}}$ is equal to 1 if the maximal cluster is non resonant and $0$ otherwise.
Note that, provided $\gamma$ is chosen to be large enough, $e^{-c_2 \gamma^{-h/\tau}\mathds 1_{\Gamma\,\mathrm{nonres}}}\le \gamma^{3h \mathds 1_{\Gamma\,\mathrm{nonres}}}$, so
\be
L(\underline l,\underline m)\le \gamma^{3h\mathds 1_{\Gamma\,\mathrm{nonres}}}
\prod_i e^{-\kappa |A_i|/2}
\prod\limits_{T\ \mathrm{n.t.}}
e^{-c_2 {M}_{T} \gamma^{-h_{T}/\tau}}.\ee

Furthermore, using the bound $e^{-\a x}\le ({\beta\over \a})^\beta e^{-\beta}x^{-\beta}$ with $\beta= 3\tau M_T$, we find
\begin{equation}
  e^{-c_2M_T \gamma^{-h_T/\tau}} \leq (\frac{c_2 e^1}{3\tau})^{-3\tau M_T} \gamma^{3M_T h_T}
  .
  \label{exp3}
\end{equation}
In addition, $\sum_{T\,\text{n.t.}} M_T \leq q$, since the clusters are nested in each other and for two clusters to be different they must differ by at least one vertex.
Now, let us introduce $M_T^\tau$ as the number of maximal non-resonant trivial clusters (i.e. maximal $\tau$-vertices) contained in $T$, and use the trivial bound $3M_T \le 2M_T+M_T^\tau$ along with (\ref{exp3}) to obtain
\begin{equation}
\prod_{T\ \mathrm{n.t.}}
e^{-c_2 M_{T} \gamma^{-h_{T}}/\tau} \le  C_3^q
.\prod_{T\ \mathrm{n.t.}} \gamma^{h_{T} (2M_{T}+M_T^\tau)}
\label{asxq}
\end{equation}
Thus, plugging this into (\ref{lap1}), we find
\bea
&&\g^h \sum^*_{\underline h,\atop \underline l, \underline m}
[L]^{1\over 2} |\l|^q (CC_3)^q
\gamma^{h\mathds 1_{\Gamma\,\mathrm{res}}}
\gamma^{3h\mathds 1_{\Gamma\,\mathrm{nonres}}}
\nn\\&&
[\prod_{T\ \mathrm{res}}
\gamma^{(h_{T}^{\text{ext}}-h_{T})}]
\prod_{T\, \text{n.t.}} \g^{h_T (2M_T+M^\tau_T)}
.
\eea
In addition,
\begin{equation}
  \prod_{T\ \mathrm{n.t.}} \gamma^{2h_T M_T}
  =
  \gamma^{-2h \mathds 1_{\Gamma\,\mathrm{nonres}}}\prod_{T\ \mathrm{nonres}}\gamma^{2h_T^{\mathrm{ext}}}
\end{equation}

\be
\g^h \sum^*_{\underline h,\atop \underline l, \underline m} [L]^{1\over 2} |\l|^q (CC_3)^q
[\prod_{T\, \text{n.t.}}
\gamma^{(h_{T}^{\text{ext}}-h_{T})}]
\prod_{T\,\text{n.t.}} \gamma^{h_{T} M^\tau_{T}} \label{ssa}
\ee
The crucial point is that the sum over the scales $h$ can be performed
summing over all the differences, taking into account that the scale $h$ is fixed.
Finally the sum over the $l,m$ is done using the factor $[L(\underline l, \underline m)]^{1\over 2}$.
(The gain term $\g^{h_T M_T^\tau}$ is dropped, as it does not lead to any significant gain.)
In conclusion the bound on a graph with $q$ vertices is $C_4^q\g^h |\l|^q$ assuming that $|\n_h|,|Z_h-1|,|v_h-1|,|w_h-1| \le C |\l|$.

\subsection{Beta function and Schwinger functions}

We are left with checking our assumption on $Z_h, \nu_h, w_h,v_h$.
We know that 
 $v_{i,\omega,h}= v_{i,\omega,h+1}-i z_{h,\o,a,b,i,1}$
with $z_{h,\o,a,b,i,1}$ expressed by the sum of renormalized Feynman graphs $\G$ such that the maximal scale of the clusters is $h+1$, an extra derivative is applied (which costs a factor $\gamma^{-h}$)
and the momenta of the external lines is fixed equal to $p_F^\o$.
Moreover by the compact support of the propagator there is at least a $\t$ vertex, as the $k=0$ value of a graph wih only $\n$ vertice is zero; therefore the
analogue of (\ref{ssa}) becomes
\be
\sum^*_{\underline h,\atop \underline l, \underline m} [L]^{1\over 2} |\l|^q (CC_3)^q    [\prod_{T\, \text{n.t.}}
\gamma^{(h_{T}^{\text{ext}}-h_{T})}]
\gamma^{2h_{T^*}}
\label{weightzz}
\ee
where $T^*$ is the non trivial cluster containing a $\t$ vertex whose scale is the largest possible (we now use the gain $\g^{h_T M_T^\tau}$ 
dropped in the bound \pref{ssa}).
In addition, summing the differences $h_T^{\mathrm{ext}}-h_T$ along a sequence of clusters that goes from $h$ to $h_{T^*}$ and discarding the others, we bound
\begin{equation}
  \sum_{T}(h_T^{\mathrm{ext}}-h_T) \leqslant h-h_{T^*}
\end{equation}
and so (\ref{weightzz}) is bounded by
\be \label{weightzz1}\g^{h\over 2}
\sum^*_{\underline h,\atop \underline l, \underline m} [L]^{1\over 2} |\l|^q (CC_3)^q    \prod_{T\, \text{n.t.}}
\gamma^{ {1\over 2}(h_{T}^{\text{ext}}-h_{T})}
.
\ee
Estimating the sum as above, we find that $|z_{h,\o,a,b,i,1}|\le C_5 \l  \g^{h\over 2}$, and $v_{i,\omega,h}= v_{i,\omega,0}-i \sum_{h'} z_{h',\o,a,b,i,1}$ hence  $v_{i,\omega,h}= v_{i,\omega,0}+O(\l)$; moreover the limiting value is reached exponentially fast 
$v_{i,\omega,h}=v_{i,\omega,-\infty}+O(\l \g^{h/2})$.
A similar argument holds for $Z_h, w_h$. Not

It remain to discuss the flow of $\n_h$; we can write, see  
\pref{ai},
\be W_{i,i,0,\alpha,\alpha'}^{h,\omega,\omega}(0,p_{F,i}^\omega)
=\g^{h+1}\n_{h+1}+ \tilde W_{i,i,0,\alpha,\alpha'}^{h,\omega,\omega}(0,p_{F,i}^\omega)\ee where $\tilde W$ is given by the sum of the terms with a number of vertices greater or equal to $2$; therefore 
\be
\nu_{h,\omega,\alpha,\alpha',i}=\g \nu_{h+1,\omega,\alpha,\alpha',i}+\b^h_\n
\label{appo2}
\ee
with $\b^h_\n=\g^{-h}
\tilde W_{i,i,0,\alpha,\alpha'}^{h,\omega,\omega}(0,p_{F,i}^\omega)$ 
is given by the sum with $q\ge 2$ of terms bounded by \pref{weightzz1}. 
We have to prove that it is possible to choose the counterterms $\n_{\omega,\alpha,\alpha',i}$ so that  $\n_{h,\omega,\alpha,\alpha',i}$ is bounded by $C \l$ for any scale $h$. Indeed from \pref{appo2} we get, $h\le -1$ 
\be
\nu_{h,\omega,\alpha,\alpha',i}=\g^{-h}(\nu_{\omega,\alpha,\alpha',i}+
\sum_{i=h}^{-1} \g^{i}\b^i_\n)
\label{appo3}
\ee
and choosing $\nu_{\omega,\alpha,\alpha',i}$ so that $\nu_{-\infty,\omega,\alpha,\alpha',i}=0$ we get
\be
\nu_{h,\omega,\alpha,\alpha',i}=-\g^{-h}\sum_{i=-\infty}^{h} \g^{i}\b^i_\n
\label{appo4}
\ee
and by using a fixed point argument we can show that there is a sequence such that
$|\nu_{h,\omega,\alpha,\alpha',i}|\le C \l\g^{h\over 2}$.

The application of the above bounds to the 2-point function, in order to derive
\pref{43}, is straightforward. The 2-point function can be written as 
\be
\hat S_{j,j}(\mathbf k+\mathbf p_{F,j}^\o )) =\sum_{h=-\infty}^0 [\hat g^{(h)}((\mathbf k+\mathbf p_{F,j}^\o ))+r^h(\mathbf k) 
\ee
where $r^h(\mathbf k) $ includes the contribution of term withs at least a vertex. We can replace in $g^{(h)}((\mathbf k+\mathbf p_{F,j}^\o ))$ the 
$v_{i,\omega,h},w_{i,\omega,h},Z_{i,\omega,h}$  with
$v_{i,\omega,-\infty},w_{i,\omega,-\infty},Z_{i,\omega,-\infty}$ obtaining the dominant term in 
\pref{43}; the subdominant term is obtained both from the 
term containing the difference betwen 
$v_{i,\omega,h}-v_{i,\omega,-\infty}$, $z_{i,\omega,h}-w_{i,\omega,-\infty}$,
$Z_{i,\omega,h}-Z_{i,\omega,-\infty}$, which have an extra factor $O(\l \g^{h/2})$, or the 
terms with at least a vertex 
which have at least
a $\n$ or a non resonant trivial vertex, with an extra $O(\g^{h/2})$
from the bounds after \pref{weightzz}.

\section{Conclusion } 

Theoretical analyses of TBG are based on the assumption of the stability of the Weyl semimetallic phase, leading to the formulation of continuum effective models.
However in lattice TBG models wth generic angles there is an emerging quasi-periodicity manifesting themself in large momenta Umklapp interactions 
that almost connect
the Dirac points, similar to the ones appearing in
electronic systems with quasi-periodic potential. 
Such terms are neglected 
in the continuum semimetallic approximations. 

In this paper we have rigorously established the stability of the semimetallic phase in a lattice model, taking into full account the large momenta Umklapp interactions.
The analysis is based on number theoretical properties of irrationals 
combined with a Renormalization Group analysis, and requires that the interlayer hopping is weak and short ranged and the the angles are chosen in a large measure set. The effect of the interaction is to produce a finite renormalization of the Dirac points and velocities. Non perturbative effects are excluded as the series are shown to be 
convergent. Compared to the Aubry-Andr\'e or similar models, the number theoretical analysis is much more 
involved due to the peculiar structure of the small divisors. 

The stability of the Weyl phases provides a justification of the use of  continnum models under the above assumptions. 
In addition, 
the present analysis paves the way to a more accurate evaluation of the 
velocities as functions of the angles, talking into account lattice or higher orders effects, and the effect 
of many body interactions, 
whose interplay with the emerging quasi-periodicity 
could lead to interesting phases. 

\begin{acknowledgements}
We thank J. Pixley for many interesting discussions.
V.M. acknowledges support from the MUR, PRIN 2022 project MaIQuFi cod. 20223J85K3.
I.J. gratefully acknowledges support through NSF Grant DMS-2349077, and the Simons Foundation, Grant Number 825876.
\end{acknowledgements}

\vfill
\pagebreak
\widetext

\appendix

\section{Fourier transform of the interlayer hopping}\label{app:fourierV}
We write $V$ in Fourier space: we get
\begin{eqnarray}
&&V=\frac \lambda{|\hat{\mathcal L}_1|^2}\sum_{x_1\in \hat{\mathcal L}_1} \sum_{x'_2\in\Lambda_2}\sum_{\alpha\in\{a,b\}}
    \int_{\mathbb R^2} \frac{dq}{4\pi^2}
    \int_{\hat{\mathcal L}_1}dk_1
    \int_{\hat{\mathcal L}_2}dk_2'\nonumber\\
&&
    e^{i(k_1x_1-k_2'x_2'+q(x_1-x_2'))}
    e^{iq(d_\alpha-Rd_{\alpha})}
    e^{i \xi (k_2'-k_1)}
    \hat\varsigma(q)\hat c^\dagger_{1,k_1,\alpha}\hat c_{2,k_2',\alpha}+\nonumber\\
&&e^{-i(k_1x_1-k_2'x_2'-q(x_1-x_2'))}
    e^{iq(d_\alpha-Rd_{\alpha})}
    e^{-i\xi(k_2'-k_1)}
    \hat\varsigma(q)\hat c^\dagger_{2,k_2',\alpha}\hat c_{1,k_1,\alpha}\label{V211}
\end{eqnarray}
and using the Poisson summation formula
\begin{equation}
  \sum_{x_1\in\mathcal L_1} e^{i (k_1+q)x_1}=|\hat{\mathcal L}_1|\sum_{l\in\mathbb Z^2} \d(k_1+q+l b)
\end{equation}
where we use the shorthand $lb\equiv l_1b_1+l_2b_2$, and
\begin{equation}
  \sum_{x_2'\in \mathcal L_2} e^{-i (k_2+q) x'_2}=|\hat{\mathcal L}_1|\sum_{m\in\mathbb Z^2} \d(k_2+q+m b')
\end{equation}
Noting that
$\hat c_{2,k_1+lb-mb',\alpha}
  \equiv
  \hat c_{2,k_1+lb,\alpha}$, $
  \hat c_{1,k_2'+mb'-lb,\alpha}
  \equiv
  \hat c_{1,k_2'+mb',\alpha}$ we finally obtain \pref{11}.
\bigskip


Rewriting (\ref{V211}) in terms of Grassmann variables, with the added imaginary time component, reads
\be\begin{array}{>\displaystyle l}
V=\frac{\beta\lambda}{|\hat\Lambda_1|^2}\sum_{x_1\in\mathcal L_1} 
\sum_{x'_2\in\mathcal L_2}\sum_{\alpha\in\{a,b\}}
\int_{\mathbb R^2} \frac{dq}{4\pi^2}
\int_{\hat\Lambda_1}d\kk_1
\int_{\hat\Lambda_2}d\kk_2'\ 
\delta(k_{1,0}-k_{2,0}')
\cdot\\[0.5cm]\cdot
\left(
e^{i(k_1x_1-k_2x_2'+q(x_1-x_2'))}
e^{iq(d_\alpha-Rd_{\alpha})}
e^{i\xi(k_2'-k_1)}
\hat\varsigma(q)\hat\psi^+_{1,\kk_1,\alpha}\hat\psi^-_{2,\kk_2',\alpha}
\label{V2112}+\right.\\[0.5cm]\indent\left.+
e^{-i(k_1x_1-k_2x_2'-q(x_1-x_2'))}
e^{iq(d_\alpha-Rd_{\alpha})}
e^{-i\xi(k_2'-k_1)}
\hat\varsigma(q)\hat\psi^+_{2,\kk_2',\alpha}\hat\psi^-_{1,\kk_1,\alpha}
\right)
\end{array}\ee
where we use the notation $\mathbf k_1=(k_{1,0},k_1)$ and $\mathbf k_2'=(k_{2,0},k_2)$.
Again, using the Poisson formula, we find \pref{112}.


\section{Proof of Lemma \ref{lemm:dioph}}\label{sec:dioph}
  

To prove lemma \ref{lemm:dioph}, we will first prove a general result on a Diophantine condition for a generic function from $[0,2\pi)$ to $\mathbb R^2$.
We will then apply this result to $|p_{F,i}^\omega,p_{F,j}^{\omega'}+lb+mb'|$, viewed as a function of $\theta$, for the various values of $i,j,\omega,\omega'$.

\subsection{Diophantine condition from $\mathbb R$ to $\mathbb R^2$}
Let us consider an interval $[\theta_0,\theta_1]\subset[0,2\pi]$, and define, given constants $C_1>0,\tau>4$ that are fixed once and for all, two twice-differentiable functions $x:[0,2\pi)\to \mathbb R,f:[0,2\pi)\to \mathbb R^2$, and a subset $\Omega(x,f)\subset[\theta_0,\theta_1]$,
\begin{equation}\label{diophantine}
\mathcal D(x,f):=
\{\theta\in \Omega(x,f):\ \forall k\in\mathbb Z^2\setminus\{0\},
\ \forall l\in\mathbb Z,\ |x(\theta)+l
+k \cdot f(\theta)|\geqslant C_1|k|^{-\tau}\}
\end{equation}
We will show that, provided $\Omega$ is chosen appropriately, under certain conditions on $f$ and $x$, $\mathcal D$ has a large measure.
The novelty of this result is that $f$ takes values in $\mathbb R^2$, but is a function of a single variable; if $f$ were a function from $\mathbb R^n$ to $\mathbb R^n$, then the fact that $\mathcal D$ has a large measure would follow from standard arguments \cite{}.
Our result is stated for $\mathbb R^2$, but it could easily be adapted to any other dimension, provided $f$ takes a single real-valued argument.
\bigskip

In order to make our argument work, we will assume that $f'(\theta)$ (the derivative of $f$) remains inside a cone, that is, we assume that $\exists\xi\in \mathbb R^2$ with $|\xi|=1$ and $\alpha\in[0,\frac\pi4)$ such that, $\forall \theta\in [\theta_0,\theta_1]$,
\begin{equation}
  f'(\theta)\in \mathcal C_\xi(\alpha)
  :=\{y\in \mathbb R^2,\ |y\cdot\xi|>|y|\cos(\alpha)\}
  .
  \label{incone}
\end{equation}
We take the set $\Omega(x,f)$ in (\ref{diophantine}) to be
\begin{equation}
\Omega(x,f):=
\{\theta\in [\theta_0,\theta_1]:\ \forall k\in \zeta,
\ |x'(\theta)
+k \cdot f'(\theta)|\ge C_3|f'(\theta)||k|^{-\epsilon}\}
\label{Omega}
\end{equation}
where $C_3>0$ is a constant, $\epsilon\in(1,\tau-3)$, and
\begin{equation}
  \zeta:=\mathbb Z^2\setminus(\{0\}\cup\mathcal C_\xi({\textstyle\frac\pi4}))
  \label{zeta}
\end{equation}
(the reason why we choose $\Omega$ in this way will become apparent in the proof of Lemma \ref{lemma:diophantine} below).



\begin{lemma}\label{lemma:diophantine}
  If the following estimates hold:
  \begin{equation}
    \min_{\theta\in [\theta_0,\theta_1]}|f'(\theta)|>0
    ,\ 
    \min_{\theta\in [\theta_0,\theta_1]}|{\textstyle\frac{\partial}{\partial\theta} (\frac{f'(\theta)}{|f'(\theta)|}})|>0
    \label{bound_df}
  \end{equation}
  $\forall \theta\in [\theta_0,\theta_1]$,
  \begin{equation}
    \min_{0<|k|<R_1}|x'(\theta)+k\cdot f'(\theta)|-C_3|f'(\theta)||k|^{-\epsilon}>0
    \label{small_dx}
  \end{equation}
  with
  \begin{equation}
    R_1:= \frac{C_3+\frac{|x'(\theta)|}{|f'(\theta)|}}{\cos(\alpha+\frac\pi4)}
    \label{R1}
  \end{equation}
  and, for some $\eta>0$,
  \begin{equation}
    \min_{0<|k|<R_2}
    |{\textstyle\frac{\partial}{\partial \theta}(\frac{x'(\theta)}{|f'(\theta)|})}+k\cdot {\textstyle\frac \partial{\partial \theta}(\frac{f'(\theta)}{|f'(\theta)|})}|
    -\eta|k||{\textstyle\frac \partial{\partial \theta}(\frac{f'(\theta)}{|f'(\theta)|})}|\cos(\alpha+{\textstyle\frac\pi4})
    >0
    \label{small_ddx}
  \end{equation}
  with
  \begin{equation}
    R_2:=
    \eta+\frac{|{\textstyle \frac \partial{\partial \theta}(\frac{x'(\theta)}{|f'(\theta)|})}|}{|{\textstyle\frac \partial{\partial \theta}(\frac{f'(\theta)}{|f'(\theta)|})}|\cos(\alpha+{\textstyle\frac\pi4})}
    \label{R2}
  \end{equation}
  then the measure of the complement of $\mathcal D$ is bounded by
  \begin{equation}
    |[\theta_0,\theta_1]\setminus\mathcal D(x,f)|
    \le
    O(C_3)+O({\textstyle \frac{C_1}{C_3 \beta}})
  \end{equation}
  where the constants in $O(\cdot)$ depend only on $\theta_0$, $\theta_1$, $x$, $f$, $\alpha$, $\epsilon$, $\tau$, $\eta$.
  In particular, if we choose $C_3\ll \theta_1-\theta_0$ and $C_1\ll (\theta_1-\theta_0)^2$, then $\mathcal D(x,f)$ fills most of $[\theta_0,\theta_1]$.
\end{lemma}

\begin{remark}
  The conditions (\ref{small_dx}) and (\ref{small_ddx}) concern a finite number of values of $k$.
  In the applications of this lemma below, we can make both of these conditions trivial by ensuring that $R_1,R_2<1$, which reduces this finite number of values for $k$ to $0$.
\end{remark}


{\it Proof}
Let 
  $|\mathcal D_\Omega^c(x,f)|$ denote the Lebesgue measure of the complement $\Omega(x,f)\setminus\mathcal D(x,f)$.
  Let
  \begin{equation}
    g_{l,k}(\th)=|x(\theta)+l+k \cdot f(\theta)|
  \end{equation}
  in terms of which
  \begin{equation}
    |\mathcal D_\Omega^c(x,f)|=\sum_{k,l}^* \int_{-C_1|k|^{-\tau}}^{C_1|k|^{-\tau}} \frac1{g_{l,k}'}dg_{l,k}
  \end{equation}
  where $\sum^*_{k,l}$ has the constraint that $\exists \theta\in[\theta_0,\theta_1]$ such that $g_{k,l}(\theta)\in[-C_1|k|^{-\tau},C_1|k|^{-\tau}]$. Therefore 
  \be
  |\mathcal D_\Omega^c(x,f)|\le
  \sum^*_{l,k} 2C_1 {|k|^{-\tau}\over \min_{\theta\in \Omega(x,f)} |x'(\th)+k\cdot f'(\theta)|}
  .
  \ee
    In addition, the number of values of $l$ such that $g_{k,l}(\theta)\in[-C_1|k|^{-\tau},C_1|k|^{-\tau}]$ is bounded by $C_2|k|$ for some constant $C_2$ (which depends only on $\theta_0,\theta_1,x,f$), and so
  \be
  |\mathcal D_\Omega^c(x,f)|\le
  2\sum_{k\in \mathbb Z^2\setminus\{0\}} C_2 C_1 {|k|^{1-\tau}\over \min_{\theta\in \Omega(x,f)} |x'(\th)+k\cdot f'(\theta)|} 
  .
  \label{bound_Dctmp}\ee
  In order for this bound to be useful, we must obtain a good lower bound on $|x'+k\cdot f'|$.

  To do so, $\Omega$ must be chosen appropriately: we wish for $k\cdot f'$ to stay as far away from $-x'$ as possible.
  Now, it cannot avoid it entirely, as $k\cdot f'$ will cover all possible values as $k$ varies in $\mathbb Z^2\setminus\{0\}$.
  By choosing $\Omega$ as in (\ref{Omega}), we ensure that $k\cdot f'$ may only approach $-x'$ for large values of $k$.
  In doing so, we can estimate $\mathcal D_\Omega^c$: we split the sum over $\mathbb Z^2\setminus\{0\}$ into a sum over $\zeta$ and a sum over its complement $\zeta^c\equiv\mathbb Z^2\cap\mathcal C_\xi(\frac\pi4)$, and compute a bound for each case.

  If $k\in \zeta$, then, by (\ref{Omega}), for $\theta\in \Omega(x,f)$,
  \begin{equation}
    |x'(\theta)+k\cdot f'(\theta)|\ge C_3|f'(\theta)||k|^{-\epsilon}
    .
    \label{boundin}
  \end{equation}
  If, on the other hand, $k\in \zeta^c\equiv\mathbb Z^2\cap\mathcal C_\xi(\frac\pi4)$,
  \begin{equation}
    |k\cdot f'(\theta)|\ge|k||f'(\theta)|\cos(\alpha+{\textstyle\frac\pi4})
  \end{equation}
  so
  \begin{equation}
    |x'(\theta)+k\cdot f'(\theta)|\ge|k||f'(\theta)|\cos(\alpha+{\textstyle\frac\pi4})-|x'(\theta)|
    .
  \end{equation}
  We distinguish two cases once more: either
  \begin{equation}
    |k|\ge \frac{C_3+\frac{|x'(\theta)|}{|f'(\theta)|}}{\cos(\alpha+\frac\pi4)}\equiv R_1
  \end{equation}
  (see (\ref{R1})) in which case (\ref{boundin}) holds true for these $k$'s as well, or $|k|<R_1$, in which case (\ref{boundin}) holds by virtue of (\ref{small_dx}).
  All in all, whatever the value of $k$, (\ref{boundin}) holds for all $k\in \mathbb Z^2\setminus\{0\}$.

  Therefore, (\ref{bound_Dctmp}) becomes
  \be
  |\mathcal D_\Omega^c(x,f)|\le
  \frac{2C_1C_2}{C_3\min_{\theta\in [\theta_0,\theta_1]} |f'(\theta)|}\sum_{k\in \mathbb Z^2\setminus\{0\}} |k|^{1-\tau+\epsilon} 
  .
  \label{bound_D}\ee
  By (\ref{bound_df}), this sum is bounded since $\epsilon <\tau-3$.

  We are left with estimating the measure of $\Omega^c(x,f):=[\theta_0,\theta_1]\setminus \Omega(x,f)$.
  Proceeding in a similar way as for $|\mathcal D_\Omega^c(x,f)|$, we find that
  \begin{equation}
    |\Omega^c(x,f)|\le 2C_3\sum_{k\in \zeta}\frac{|k|^{-\epsilon}}{\min_{\theta\in[\theta_0,\theta_1]}|\frac{\partial}{\partial \theta}( \frac{x'(\theta)}{|f'(\theta)|})+k\cdot \frac\partial{\partial \theta}(\frac{f'(\theta)}{|f'(\theta)|})|}
    .
  \end{equation}
  Here we see why it was necessary to introduce the set $\zeta$ in (\ref{Omega}): if $\zeta$ were taken to be $\mathbb Z^2\setminus\{0\}$, then we would run into exactly the same problem as before: the denominator in this bound would not be bounded away from 0.
  However, the problem that gave rise to the necessity of introducing $\Omega$ in the first place actually only occurred for the $k$'s that are close to being orthogonal to $f'(\theta)$.
  So we can restrict $\zeta$ to only include the $k$'s that are close to being orthogonal to $f'(\theta)$, which is why we define $\zeta$ as in (\ref{zeta}).
  Because $\frac\partial{\partial \theta}( \frac{f'}{|f'|})$ is orthogonal to $f'$,
  \begin{equation}
    {\textstyle\frac\partial{\partial \theta}(\frac{f'}{|f'|})}\in \mathcal C_{\xi^\perp}(\alpha)
  \end{equation}
  and so, for $k\in \zeta$, because the maximal angle between $k$ and ${\textstyle\frac\partial{\partial \theta}(\frac{f'}{|f'|})}$ is $\alpha+\frac\pi4$,
  \begin{equation}
    |k\cdot {\textstyle\frac \partial{\partial \theta}(\frac{f'(\theta)}{|f'(\theta)|})}|>
    |k||{\textstyle\frac \partial{\partial \theta}(\frac{f'(\theta)}{|f'(\theta)|})}|\cos(\alpha+{\textstyle\frac\pi4})
    .
  \end{equation}
  Thus,
  \begin{equation}
    |{\textstyle\frac \partial{\partial \theta}(\frac{x'(\theta)}{|f'(\theta)|})}+k\cdot {\textstyle\frac \partial{\partial \theta}(\frac{f'(\theta)}{|f'(\theta)|})}|
    >
    |k||{\textstyle\frac \partial{\partial \theta}(\frac{f'(\theta)}{|f'(\theta)|})}|\cos(\alpha+{\textstyle\frac\pi4})
    -|{\textstyle\frac \partial{\partial \theta}(\frac{x'(\theta)}{|f'(\theta)|})}|
    .
  \end{equation}
  Therefore, if
  \begin{equation}
    |k|\ge
    \eta+\frac{|{\textstyle \frac \partial{\partial \theta}(\frac{x'(\theta)}{|f'(\theta)|})}|}{|{\textstyle\frac \partial{\partial \theta}(\frac{f'(\theta)}{|f'(\theta)|})}|\cos(\alpha+{\textstyle\frac\pi4})}
    \equiv R_2
  \end{equation}
  then
  \begin{equation}
    |{\textstyle\frac \partial{\partial \theta}(\frac{x'(\theta)}{|f'(\theta)|})}+k\cdot {\textstyle\frac \partial{\partial \theta}(\frac{f'(\theta)}{|f'(\theta)|})}|
    >\eta|k||{\textstyle\frac \partial{\partial \theta}(\frac{f'(\theta)}{|f'(\theta)|})}|\cos(\alpha+{\textstyle\frac\pi4})
    .
    \label{tmpineq}
  \end{equation}
  If, on the other hand, $|k|<R_2$, then (\ref{tmpineq}) holds by virtue of (\ref{small_ddx}).
  Thus, (\ref{tmpineq}) holds for all $k\in \zeta$.
  Therefore,
  \begin{equation}
    |\Omega^c(x,f)|
    \le
    \frac{2C_3}{\eta{\displaystyle\min_{\theta\in[\theta_0,\theta_1]}}|\frac\partial{\partial \theta}(\frac{f'(\theta)}{|f'(\theta)|})|\cos(\alpha+\frac\pi4)}
    \sum_{k\in \mathbb Z^2\setminus\{0\}}|k|^{-1-\epsilon}
    .
    \label{bound_Omega}
  \end{equation}
  By (\ref{bound_df}), this sum is bounded since $\epsilon>1$.
  We conclude the proof by combining (\ref{bound_D}) with (\ref{bound_Omega}).
\qed

\subsection{Applying the Diophantine condition to $|p_{F,i}^\omega-p_{F,j}^{\omega'}+lb+mb'|$}
We now apply lemma \ref{lemma:diophantine} repeatedly to prove lemma \ref{lemm:dioph}.

Let us first consider the case $i=j$, $\omega=\omega'$, and $y=m$, that is, we wish to find a condition on $\theta$ such that
\begin{equation}
  |p_{F,i}^\omega-p_{F,j}^{\omega'}+lb+mb'|
  \equiv |lb+mb'|
  \equiv |M|
  \ge\frac{C_0}{|m|^\tau}
  \label{cond11}
\end{equation}
(recall (\ref{cond}) and (\ref{M})).
We recall (\ref{Mineq}):
\begin{equation}
  |M|\ge
  \frac{2\pi}{\sqrt3}(l_\omega+m\cdot f_\omega)
\end{equation}
with $l_+\equiv l_1$ and $l_-\equiv l_2$, and
\begin{equation}
  f_\omega(\theta):=({\textstyle \frac{\varphi_1-\omega\varphi_3}2},\ {\textstyle \frac{\varphi_2+\omega\varphi_4}2})
  .
\end{equation}
Now, recalling the definition (\ref{diophantine}), we have that if $\theta\in \mathcal D(0,f_\omega)$, then the inequality (\ref{cond}) with $i=i'$, $\omega=\omega'$, and $y=m$ holds with $C_0:=\frac{2\pi}{\sqrt3}C_1$.
We therefore just need to use Lemma \ref{lemma:diophantine} to ensure that the measure of this set is large.
Taking $\theta_0,\theta_1$ sufficiently small, it suffices to verify the conditions at $\theta=0$, and conclude by continuity.
In particular, when $\theta_0,\theta_1$ are small, $f'(\theta)$ will take values in a small cone $\mathcal C_{f'(0)}(\alpha)$ with $\alpha=O(\theta_1)$.
Next, by a straightforward computation, we find
\begin{equation}
  |f'_\omega(0)|=\sqrt{\frac53}
  ,\quad
  \left|\frac\partial{\partial \theta}\frac{f'_\omega(0)}{|f'_\omega(0)|}\right|=\frac{2\sqrt3}5
\end{equation}
Which are both non-zero, so (\ref{bound_df}) is satisfied for small enough $\theta$.
Since $x=0$, the other assumptions trivially hold: we choose $C_3<1/\sqrt2$ and $\eta<1$ such that $R_1,R_2<1$, in which case the minima in (\ref{small_dx}) and (\ref{small_ddx}) are taken over empty sets, so (\ref{small_dx}) and (\ref{small_ddx}) hold trivially.
Thus, by Lemma \ref{lemma:diophantine}, choosing $C_1=O(\theta_1^2)$, the set $\mathcal D(0,f_\omega)$ has a large measure.

We now repeat the argument for the other values of $i,j$, $y$, and $\omega,\omega'$.
First, note that $p_F^+-p_F^-=\frac13(b_1-b_2)$ so the condition (\ref{cond}) holds for $\o\neq \o'$ whenever it holds for $\omega=\omega'$.
Next, note that
\begin{equation}
  |p_{F,i}^\omega-p_{F,j}^\omega+lb+mb'|
  =
  |R^T(p_{F,i}^\omega-p_{F,j}^\omega+lb+mb')|
\end{equation}
which corresponds to exchanging $m$ and $l$, and flipping the sign of $\theta$.
The arguments made for $\theta$ may be adapted in a straightforward way to the case $-\theta$ so  our derivation for $y=m$ also applies to $y=l$.

We are thus left with the case $i\neq j$, $\omega=\omega'$, and $y=m$.
Without loss of generality, we choose $i=1$, $j=2$, and we wish to bound
\begin{equation}
  |p_{F,1}^\omega-p_{F,2}^{\omega}+lb+mb'|
  \ge\frac{C_0}{|m|^\tau}
  \label{cond12}
\end{equation}
Proceeding as we did above, we bound
\begin{equation}
  |p_{F,1}^\omega-p_{F,2}^{\omega}+lb+mb'|
  \ge
  \frac{2\pi}{\sqrt3}\left(
    x_\omega(\theta)
    +l_1+m\cdot f_\omega(\theta)
  \right)
\end{equation}
with
\begin{equation}
  x_\omega(\theta):=\frac{1}3(1-c_\theta)+\omega\frac{1}{\sqrt3}s_\theta
  .
\end{equation}
Therefore, if $\theta\in \mathcal D(x_\omega,f_\omega)$, then (\ref{cond12}) holds with $C_0=\frac{2\pi}{\sqrt3}C_1$.
To show that this set has a large measure, we check the assumptions of Lemma \ref{lemma:diophantine}, as we did above.
Again, we check the assumptions at $\theta=0$, and argue by continuity.
We compute
\begin{equation}
  |x'_\omega(0)|=\frac1{\sqrt3}
  ,\quad
  \left|\frac\partial{\partial \theta}\frac{x'_\omega(0)}{|f'_\omega(0)|}\right|=\frac{8}{5\sqrt{15}}
\end{equation}
We thus find that if $C_3<1/\sqrt2-1/\sqrt5$ and $\eta<1-4\sqrt2/\sqrt{45}$, then $R_1,R_2<1$, so the minima in (\ref{small_dx}) and (\ref{small_ddx}) are taken over empty sets, so (\ref{small_dx}) and (\ref{small_ddx}) hold trivially.
\bigskip

All in all, we have found that if we restrict the values of $\theta$ to an intersection of Diophantine sets:
\begin{equation}
  \theta\in
  \bigcap_{\omega=\pm}\bigcap_{\sigma=\pm}\mathcal D(0,f_\omega(\sigma \theta))
  \cap
  \bigcap_{\omega=\pm}\bigcap_{\sigma=\pm}\mathcal D(x_\omega(\sigma \theta),f_\omega(\sigma \theta))
\end{equation}
then (\ref{cond}) is satisfied for any value of $i,i'$, $\omega,\omega'$, and $y$ with a constant $C_0=O(\theta_1^2)$.
Because each set has an arbitrarily large measure (relative to $[\theta_0,\theta_1]$), their intersection also does.






\section{Naive perturbation theory} \label{app:explicit_feynman}

The Schwinger function is computed using perturbation theory: formally,
\begin{equation}
\begin{array}{>\displaystyle l}
  (S_{1,1}(\mathbf k))_{\alpha',\alpha}=
  \sum_{N=0}^\infty
  \sum_{\alpha_0,\cdots,\alpha_{2N+1}}
  (g_1(\mathbf k))_{\alpha,\alpha_0}
  \left(\prod_{n=0}^N
  \left(
    \sum_{l_{2n}}\tau_{l_{2n},\alpha_{2n}}^{(1)}(k+m_{2n-1}b'+l_{2n}b)
    (g_2(\mathbf k+l_{2n}b))_{\alpha_{2n},\alpha_{2n+1}}
    \cdot\right.\right.\\\hfill\cdot\left.\left.
    \sum_{m_{2n+1}}\tau_{m_{2n+1},\alpha_{2n+1}}^{(2)}(k+l_{2n}b+m_{2n+1}b')
    ((g_1(\mathbf k+m_{2n+1}b'))_{\alpha_{2n+1},\alpha_{2n+2}})^{\mathds 1_{n<N}}
  \right)\right)
  (g_1(\mathbf k))_{\alpha_{2N+1},\alpha'}
  \label{linear1}
\end{array}\end{equation}
where $m_{-1}\equiv l_{-1}\equiv0$ and $\mathds 1_{n<N}\in\{0,1\}$ is equal to 1 if and only if $n<N$,
\begin{equation}\begin{array}{>\displaystyle l}
  (S_{2,2}(\mathbf k))_{\alpha',\alpha}=
  \sum_{N=0}^\infty
  \sum_{\alpha_0,\cdots,\alpha_{2N+1}}
  (g_2(\mathbf k))_{\alpha,\alpha_0}
  \left(\prod_{n=0}^N
  \left(
    \sum_{m_{2n}}\tau_{m_{2n},\alpha_{2n}}^{(2)}(k+l_{2n-1}b+m_{2n}b')
    (g_1(\mathbf k+m_{2n}b'))_{\alpha_{2n},\alpha_{2n+1}}
    \cdot\right.\right.\\\hfill\cdot\left.\left.
    \sum_{l_{2n+1}}\tau_{l_{2n+1},\alpha_{2n+1}}^{(1)}(k+m_{2n}b'+l_{2n+1}b)
    ((g_2(\mathbf k+l_{2n+1}b))_{\alpha_{2n+1},\alpha_{2n+2}})^{\mathds 1_{n<N}}
  \right)\right)
  (g_2(\mathbf k))_{\alpha_{2N+1},\alpha'}
\end{array}\end{equation}
\begin{equation}\begin{array}{>\displaystyle l}
  (S_{2,1}(\mathbf k))_{\alpha',\alpha}=
  \sum_{N=0}^\infty
  \sum_{\alpha_0,\cdots,\alpha_{2N}}
  (g_1(\mathbf k))_{\alpha,\alpha_0}
  \left(\prod_{n=0}^N
  \left(
    \sum_{l_{2n}}\tau_{l_{2n},\alpha_{2n}}^{(1)}(k+m_{2n-1}b'+l_{2n}b)
    ((g_2(\mathbf k+l_{2n}b))_{\alpha_{2n},\alpha_{2n+1}})^{\mathds 1_{n<N}}
    \cdot\right.\right.\\\hfill\cdot\left.\left.
    \left(\sum_{m_{2n+1}}\tau_{m_{2n+1},\alpha_{2n+1}}^{(2)}(k+l_{2n}b+m_{2n+1}b')
    (g_1(\mathbf k+m_{2n+1}b'))_{\alpha_{2n+1,2n+2}}\right)^{\mathds 1_{n<N}}
  \right)\right)
  (g_2(\mathbf k))_{\alpha_{2N,\alpha'}}
\end{array}\end{equation}
\begin{equation}\begin{array}{>\displaystyle l}
  (S_{1,2}(\mathbf k))_{\alpha',\alpha}=
  \sum_{N=0}^\infty
  \sum_{\alpha_0,\cdots,\alpha_{2N}}
  (g_2(\mathbf k))_{\alpha,\alpha_0}
  \left(\prod_{n=0}^N
  \left(
    \sum_{m_{2n}}\tau_{m_{2n},\alpha_{2n}}^{(2)}(k+l_{2n-1}b+m_{2n}b')
    ((g_1(\mathbf k+m_{2n}b'))_{\alpha_{2n},\alpha_{2n+1}})^{\mathds 1_{n<N}}
    \cdot\right.\right.\\\hfill\cdot\left.\left.
    \left(\sum_{l_{2n+1}}\tau_{l_{2n+1}}^{(1)}(k+m_{2n}b'+l_{2n+1}b)
    (g_2(\mathbf k+l_{2n+1}b))_{\alpha_{2n+1},\alpha_{2n+2}}\right)^{\mathds 1_{n<N}}
  \right)\right)
  \label{linear4}
\end{array}\end{equation}


\section {Symmetry constraints on the resonant terms}\label{sec:symm}

\subsection{Symmetries of the system}\label{sec:symmetry}

Let us state the symmetries of the model, which will play an important role in our discussion.
Note that the $C_2T$ symmetry in (c) only holds if $\xi=(0,1/2)$.
\vskip.3cm
a) {\it  Complex conjugation}:
every complex constant is conjugated and
\be
\hat\psi_{1,k_0,k,\alpha}^\pm\mapsto
  e^{\mp i \xi(b_1+b_2)}
  \hat\psi_{1,-k_0, -k,\alpha}^\pm
  ,\quad
  \label{sim}
\hat\psi_{2,k_0,k,\alpha}^\pm\mapsto 
  e^{\mp i \xi(b'_1+b'_2)}
  \hat\psi_{2,-k_0,-k,\alpha}^\pm\ee
Indeed, it is straighforward to check that $H_1$ and $H_2$ (see (\ref{H1k}) and (\ref{H2k})) are left invariant by (\ref{sim}) (following from the fact that $\Omega(-k)=\Omega(k)^*$).
To check the invariance of the interlayer hopping term $V$ (see (\ref{V2112})), there is one subtelty: because $\hat{\mathcal L}_1$ and $\hat{\mathcal L}_2$ have different periodicity, we cannot simply change $k_1$ to $-k_1$ and $k_2'$ to $-k_2'$, which would not leave $\hat{\mathcal L}_i$ invariant.
Instead, we map $k_1$ to $-k_1+b_1+b_2$ and $k_2'$ to $-k_2'+b_1'+b_2'$.
It is then straightforward to check (using (\ref{V2112})) that $V$ remains invariant under (\ref{sim}), using $e^{i(1,1)bx_1}=e^{i(1,1)b'x_2'}=1$
and periodicity.
\vskip.3cm

b)
 {\it Particle-hole} \be
\hat\psi_{1,k_0, k,\alpha}^\pm\mapsto
  ie^{\pm i\xi(b_1+b_2)}
  \hat\psi_{1,k_0,-k,\alpha}^\mp
  ,\quad
  \label{sim1}
  \hat\psi_{2,k_0, k,\alpha}^\pm\mapsto 
  ie^{\pm i\xi(b'_1+b'_2)}
  \hat\psi_{2,k_0, -k,\alpha}^\mp\ee
The argument is substantially the same as for the conjugation symmetry (a).
\vskip.3cm

 c) {\it $C_2 T$ symmetry}
\be
\hat\psi_{j,k_0,k,\alpha}^\pm\mapsto
  e^{\pm i\chi_j(k)}
  \hat\psi_{j,k_0,-k,\bar\alpha}^\pm
  \label{C2T}
\ee
where if $\alpha=a$ then $\bar \alpha=b$ and if $\alpha=b$ then $\bar \alpha =a$, and
\begin{equation}
  \chi_1(k):=
  \frac{d_b}2(b_1+b_2)-kd_b-\sigma_{k,2}bd_b
  ,\quad
  \chi_2(k):=
  \frac{d_b}2(b_1'+b_2')-kRd_b-\sigma_{k,1}b'd_b
  .
\end{equation}
$H_1$ and $H_2$ are invariant under (\ref{C2T}) for the same reason as the conjugation and particle-hole symmetries.
For the interlayer hopping, it is easiest to use the expression of $V$ in (\ref{11}), which involves two integrals: one over $\hat{\mathcal L}_1$ and one over $\hat{\mathcal L}_2$.
Let us discuss the invariance of the integral over $\hat{\mathcal L}_1$, as the invariance of the other follows from a similar argument.
We change variables in the integral: $k\mapsto -k+b_1+b_2$, as well as in the sum over $l$: $(l_1,l_2)\mapsto -(l_1+1,l_2+1)$, and in the sum over $\alpha$: $\alpha\mapsto\bar \alpha$.
This changes $\tau_{l,\alpha}^{(1)}(k+lb)$ to $\tau_{-l-(1,1),\bar \alpha}^{(1)}(-k-lb)$.
Now, by (\ref{tau1}),
\begin{equation}
  \frac{\tau_{-l-(1,1),\bar \alpha}^{(1)}(-k-lb)}{\tau_{l,\alpha}^{(1)}(k+lb)}
  =
  e^{-i\xi(2l+(1,1))b}
  e^{i(k+lb)(d_{\bar \alpha}+d_\alpha-Rd_{\alpha}-Rd_{\bar \alpha})}
  e^{-i\xi (\sigma_{-k-lb,1}-\sigma_{k+lb,1})b'}
  \frac{\hat\varsigma(k+lb)}{\hat\varsigma^*(k+lb)}
  .
\end{equation}
In addition, if $k+lb-\sigma_{k+lb,1}b'\in\hat{\mathcal L}_2$, then $-k-lb+(\sigma_{k+lb,1}+(1,1))b'\in\hat{\mathcal L}_2$, so
\begin{equation}
  \sigma_{-k-lb,1}=-\sigma_{k+lb,1}-(1,1)
\end{equation}
and, since $d_a=0$ and $d_b=(1,0)$,
\begin{equation}
  d_{\bar \alpha}+d_\alpha-Rd_{\alpha}-Rd_{\bar \alpha}
  =d_b-Rd_b
\end{equation}
and, since $\varsigma(x)=\varsigma(-x)$, $\hat\varsigma\in \mathbb R$.
Therefore,
\begin{equation}
  \frac{\tau_{-l-(1,1),\bar \alpha}^{(1)}(-k-lb)}{\tau_{l,\alpha}^{(1)}(k+lb)}
  =
  e^{-i\xi(2l+(1,1))b}
  e^{i(k+lb)(d_b-Rd_b)}
  e^{i\xi (2 \sigma_{k+lb,1}+(1,1))b'}
  .
\end{equation}
Since $\xi=d_b/2$,
\begin{equation}
  \frac{\tau_{-l-(1,1),\bar \alpha}^{(1)}(-k-lb)}{\tau_{l,\alpha}^{(1)}(k+lb)}
  =
  e^{ikd_b}
  e^{-i(k+lb)Rd_b}
  e^{i d_b\sigma_{k+lb,1}b'}
  e^{i\frac{d_b}2(b_1'+b_2'-b_1-b_2)}
  .
\end{equation}
It is then straightforward to check that this extra phase gets canceled out exactly by $e^{\pm i\chi_j}$ in (\ref{C2T}) (to see this, note that if $k\in\hat{\mathcal L}_1$, then $\sigma_{k,2}=0$).


\vskip.3cm

d) {\it Inversion} 
\be \hat\psi_{j,k_0,k,\alpha}^\pm\to i (-1)^\alpha (-1)^j\hat\psi_{j,-k_0,k,\alpha}^\pm\label{inversion}\ee
\vskip.3cm 
It is straightforward to check that $H_1$, $H_2$, and the interlayer hopping (using (\ref{V2112})) are invariant under (\ref{inversion}).

\subsection {Constraints on the resonant terms}

The discrete symmetry properties seen above implies
severely constraint the form of the resonant terms. 
In the following, we use the notation ``$=^a$'' to mean ``by using symmetry (a) from Section \ref{sec:symmetry} (that is, Complex conjugation), it is equal to'', and similarly for ``$=^b$'', ``$=^c$'', ``$=^d$''.

\begin{enumerate}

\item
Using that $W_{aa}(k_0,k)=^d-W_{aa}(-k_0,k)$
we get  $W_{aa}(0,p_F^\omega)=\partial_1W_{aa}(0,p_F^\omega)=\partial_2W_{aa}(0,p_F^\omega)=0$.
Similarly,
$W_{bb}(0,p_F^\omega)=\partial_1W_{bb}(0,p_F^\omega)=\partial_2W_{bb}(0,p_F^\omega)=0$.
\item
From $W_{ab}(k_0,k)=^b W_{ba}(k_0,-k)=^a W^*_{ba}(-k_0,k)$
we get $W_{ab}(0,p_F^\omega)=W^*_{ba}(0,p_F^\omega)$,
$\partial_1W_{ab}(0,p_F^\omega)=\partial_1W^*_{ba}(0,p_F^\omega)$,
$\partial_2W_{ab}(0,p_F^\omega)=\partial_2W^*_{ba}(0,p_F^\omega)$.
Moreover $W_{ab}(k_0,k)=^d W_{ab}(-k_0,k)$ hence $\partial_0 W_{ab}(0,p_F^\omega)=0$.

\item
$\partial_0 W_{aa}(k_0,k)=^a-\partial_0 W_{aa}^*(-k_0,-k)=^b-\partial_0 W_{aa}^*(-k_0,k)$
hence is pure imaginary at $k_0=0$; moreover
$\partial_0 W_{aa}(k_0,k)=^c \partial_0 W_{bb}(k_0,-k)=^a-\partial_0 W^*_{bb}(-k_0,k)$
so that $\partial_0 W_{aa}(0,p_F^\o)=\partial W_{bb}(0,p_F^\o)=i z$ with $z$ real

\end{enumerate}



\vfill
\eject


\bibliographystyle{amsalpha}
\begin{thebibliography}{19}


\bibitem{a}
Y. Cao, V. Fatemi, A. Demir, S. Fang, S. L. Tomarken, J. Y. Luo, J. D. Sanchez-Yamagishi, K. Watanabe, T. Taniguchi, E. Kaxiras, R. C. Ashoori, and P. Jarillo-Herrero. Correlated insulator behaviour at half-filling in magic-angle graphene superlattices.
Nature 556, 80 (2018).

\bibitem{b}
Eve Andrei \& Allan H. MacDonald. Graphene Bilayers with a Twist
Nature Materials 19, 1265 (2020)

\bibitem{1} Lopes dos Santos JMB, Peres NMR, Castro Neto AH. 
Graphene bilayer with a twist: electronic structure.
Phys. Rev. Lett. 99, 256802 (2007)


\bibitem{2}
R. Bistritzer and A. H. MacDonald. Transport between twisted graphene
layers. Phys. Rev. B . 81, 24, 1.
(2010)

\bibitem{3}R. Bistritzer and A. H. MacDonald. Moir´e bands in twisted double-layer
graphene. Proceedings of the National Academy of Sciences of the United
States of America 108.30, 1223 (2011)

\bibitem{1aa1}
i Chun Po, Liujun Zou, Ashvin Vishwanath, and T. Senthil
Origin of Mott insulating behavior and superconductivity in twisted bilayer graphene.
Phys. Rev. X 8, 031089 (2018)

\bibitem{1aa2}
Liujun Zou, Hoi Chun Po, Ashvin Vishwanath, and T. Senthil. Band structure of twisted bilayer graphene: Emergent symmetries, commensurate approximants, and wannier obstructions. Phys. Rev. B, 98(8), 5435.  (2018)

\bibitem{1bb} S. Shallcross, S. Sharma, E. Kandelaki, and
O. A. Pankratov. Electronic structure of turbostratic
graphene, Phys. Rev. B 81, 165105 (2010).

\bibitem{1aaa}
G. Trambly de Laissardi`ere, D. Mayou, and L. Magaud, Localization of dirac electrons in rotated
graphene bilayers. Nano Letters 10, 804–808 (2010)


\bibitem{H1} 
Huaqing Huang, Yong-Shi Wu, and Feng Liu. Aperiodic topological crystalline insulators.
Phys. Rev. B 101, 041103(R). ( 2020)

\bibitem{H2}T. Cea
, Pierre A. Pantaleón, and Francisco Guine
Band structure of twisted bilayer graphene on hexagonal boron nitride
Phys. Rev. B 102, 155136 (2020)

\bibitem{H3}Jingtian Shi
, Jihang Zhu, and A. H. MacDonald
Moiré commensurability and the quantum anomalous Hall effect in twisted bilayer graphene on hexagonal boron nitride.
Phys. Rev. B 103, 075122. (2021)

\bibitem{H4} Dan Mao, T. Senthil
Quasiperiodicity, band topology, and moiré graphene
Phys. Rev. B 103, 115110 (2021)

\bibitem{H5} Michael G. Scheer
, Kaiyuan Gu
, and Biao Lian
Magic angles in twisted bilayer graphene near commensuration: Towards a hypermagic regime.
Phys. Rev. B 106, 115418 (2022)


\bibitem{P0}
Yixing Fu, Elio J. König, Justin H. Wilson, Yang-Zhi Chou \& Jedediah H. Pixley. 
Magic-angle semimetals
npj Quantum Materials 5, 71. (2020) 

\bibitem{P1}
Jinjing Yi, Elio J. König2,J. H. Pixley
Low energy excitation spectrum of magic-angle semimetals
Phys. Rev. B 106, 195123 (2022)

\bibitem{P2}
H Pixley, DA Huse, JH Wilson
Connecting the avoided quantum critical point to the magic-angle transition in three-dimensional Weyl semimetals
Physical Review B, 109, 165151. (2024 )



\bibitem{Au} S. Aubry and G. André, Ann. Israel Phys. Soc. 3, 133 (1980).

\bibitem{DS} E. Dinaburg, E, Y. Sinai, Funct. an. and its app. 9, 279
(1975)

\bibitem{FS}J. Fröhlich
T. Spencer, P. Wittwer
Localization for a class of one-dimensional quasi-periodic Schrödinger operators. 
Comm. Math. Phys. 132(1) (1990)


\bibitem{Q0}J Vidal,
D Mouhanna,
T Giamarchi.
Interacting fermions in self-similar potentials.
Phys. Rev. B 65, 014201 (2001)

\bibitem{M10}G. Benfatto, G.Gentile, V. Mastropietro.
Electrons in a lattice with an incommensurate potential. J. Stat. Phys.
89, pages 655–708, (1997)

\bibitem{Q1}V. Mastropietro. Small Denominators and Anomalous Behaviour in the Incommensurate Hubbard–Holstein Model
 Commun. Math. Phys. 201, 81 (1999);Dense gaps and scaling relations in the interacting Aubry-Andre' model Phys. Rev. B 93, 245154 (2016)

\bibitem{Q2}
V. Oganesyan and D. A. Huse. Localization of interacting fermions at high temperature.
 Phys. Rev. B 75, 155111 (2007).

\bibitem{Q3}V. Mastropietro
Localization of Interacting Fermions in the Aubry-André Model
Phys. Rev. Lett. 115, 180401 (2015)

\bibitem{M11}  V. Mastropietro. Localization of Interacting Fermions in the Aubry-André Model
Phys. Rev. Lett. 115, 180401 (2015); Localization in Interacting Fermionic Chains with Quasi-Random Disorder,
Comm.Math. Phys.
Volume 351, pages 283–309, (2017)

\bibitem{P4}
J. H. Pixley, Justin H. Wilson, David A. Huse, Sarang
Gopalakrishnan 
Weyl Semimetal to Metal Phase Transitions Driven by Quasiperiodic Potentials Phys. Rev. Lett. 120, 207604 (2018)

\bibitem{W0}
J. H. Pixley, J. H. Wilson, D. A. Huse, and S. Gopalakrishnan,
Weyl Semimetal to Metal Phase Transitions Driven by Quasiperiodic Potentials.
 Phys. Rev. Lett. 120, 207604 (2018).

\bibitem{W1}V. Mastropietro
Stability of Weyl semimetals with quasiperiodic disorder.
Phys. Rev. B 102, 045101 (2020)



\bibitem{Lu} J.M. Luck, Critical behavior of the aperiodic quantum Ising chain in a transverse magnetic field, J.
Stat. Phys., 72, 417  (1993)

\bibitem{Ca} P. Crowley, A. Chandran, and C. Laumann, Quasiperiodic quantum Ising transitions in 1d. Physical
Review Letters, 120 (2018).

\bibitem{Ga} M. Gallone, V. Mastropietro. Universality in the 2d Quasi-periodic Ising Model and Harris–Luck Irrelevance.
Commun. Math. Phys. 405, 235 (2024)

\bibitem{111}A. B. Watson, Tianyu Kong, Allan H. MacDonald , Mitchell Luskin 
Bistritzer–MacDonald dynamics in twisted bilayer graphene.
J. Math. Phys. 64, 031502 (2023)



\end{thebibliography}
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