Peierls substitution

The Peierls substitution method, named after the original work by R. Peierls ^[1] is a widely employed approximation for describing tightly-bound electrons in the presence of a slowly varying magnetic vector potential. ^[2]

In the presence of an external vector potential $\mathbf{A}$ the translation operators, which form the kinetic part of the Hamiltonian in the tight-binding framework, are simply -

${\displaystyle \mathbf{T}_x = |m+1,n\rangle\langle m,n|e^{i\theta^x_{m,n}}, \quad \mathbf{T}_y = |m,n+1\rangle\langle m,n|e^{i\theta^y_{m,n}} }$

and in the second quantization formulation

${\displaystyle \mathbf{T}_x = \boldsymbol{\psi}^\dagger_{m+1,n}\boldsymbol{\psi}_{m,n}e^{i\theta^x_{m,n}}, \quad \mathbf{T}_y = \boldsymbol{\psi}^\dagger_{m,n+1}\boldsymbol{\psi}_{m,n}e^{i\theta^y_{m,n}}. }$

The phase factors are defined as

${\displaystyle \theta^x_{m,n}=\frac{e}{\hbar}\int_m^{m+1} A_x(x,n)\text{d}x, \quad \theta^y_{m,n}=\frac{e}{\hbar}\int_n^{n+1} A_y(m,y) \text{d}y. }$

Properties of the Peierls substitution

1. The number of flux quanta per plaquette $\phi_{mn}$ is related to the lattice curl of the phase factor,

${\displaystyle \begin{align} \boldsymbol{\nabla}\times\theta_{m,n}&=\Delta_x\theta^y_{m,n}-\Delta_y\theta^x_{m,n}=\left(\theta^y_{m+1,n}-\theta^y_{m,n}-\theta^x_{m,n+1}+\theta^x_{m,n}\right)\\ &=\frac{e}{\hbar}\int_{\text{unit cell}}\mathbf{A}\cdot \text{d}\mathbf{l}=2\pi\frac{e}{h}\int \mathbf{B} \cdot \text{d}\mathbf{s}=2\pi\phi_{m,n} \end{align} }$ and the total flux through the lattice is ${\textstyle \Phi=\Phi_0\sum_{m,n}\phi_{m,n} }$ with $\Phi_0=hc/e$ in cgs units (and $\Phi_0=2\pi$ in natural units).

2. flux quanta per plaquette $\phi_{mn}$ is related to the accumulated phase of a single particle state, $|\psi\rangle=\boldsymbol{\psi}_{i,j}|0\rangle$ surrounding a plaquette:

${\displaystyle \begin{align} \mathbf{T}_y^\dagger \mathbf{T}_x^\dagger \mathbf{T}_y\mathbf{T}_x|\psi\rangle&=\mathbf{T}_y^\dagger \mathbf{T}_x^\dagger \mathbf{T}_y |i+1,j\rangle e^{i\theta^x_{i,j}}= \mathbf{T}_y^\dagger \mathbf{T}_x^\dagger |i+1,j+1\rangle e^{i\left( \theta^x_{i,j}+\theta^y_{i+1,j} \right)}\\ &=\mathbf{T}_y^\dagger |i,j+1\rangle e^{i\left( \theta^x_{i,j}+\theta^y_{i+1,j}-\theta^x_{i,j+1} \right)}= |i,j\rangle e^{i\left( \theta^x_{i,j}+\theta^y_{i+1,j}-\theta^x_{i,j+1}-\theta^y_{i,j} \right)}=|i,j\rangle e^{i2\pi \phi_{m,n}} \end{align} }$

Justification of Peierls substitution

Here we give three derivations of the Peierls substitution, each one is based on a different formulation of quantum mechanics theory.

The axiomatic approach

Here we give a simple derivation of the Peierls substitution, which is based on the Feynman's Lectures (Vol. III, Chapter 21)^[3] . This derivation postulate that magnetic fields are incorporated in the tight-binding model by adding a phase to the hopping terms and show that it is consistent with the continuum Hamiltonian. Thus, our starting point is the Hofstader Hamiltonain:^[2]

${\displaystyle H_0=\sum_{m,n}\bigg(-te^{i\theta^x_{m,n}}\vert m\!+\!a,n \rangle \langle m,n\vert -te^{i\theta_{m,n}^y}\vert m,n\!+\!a\rangle\langle m,n\vert -\epsilon_0\vert m,n\rangle\langle m,n\vert\bigg)+ \text{h.c}. }$

The translation operator $\vert m+1\rangle\langle m\vert$ can be written explicitly using its generator, that is the momentum operator. Under this representation its easy to expand it up to the second order,

${\displaystyle \vert m\!+\!a\rangle\langle m\vert= \exp{\bigg(\!-\!\frac{\mathbf{p}_xa}{\hbar}\bigg)}\vert m\rangle\langle m\vert =\left(1-\frac{i\mathbf{p}_x}{\hbar}a -\frac{\mathbf{p}_x^2}{2\hbar^2}a^2+\mathcal{O}(a^3) \right)\vert m\rangle\langle m\vert }$

and in a 2D lattice $\vert m\!+\!a\rangle\langle m\vert \longrightarrow\vert m\!+\!a,n\rangle\langle m,n\vert$. Next, we expand up to the second order the phase factors,

${\displaystyle e^{i\theta}=1-i\theta^\prime a-\frac{1}{2}{\theta^\prime}^2a^2-\frac{i}{2}{\theta^{\prime\prime}}^2a^2=1+\frac{ieA_x}{h}a-\frac{e^2A_x^2}{2\hbar^2}a^2+\frac{ieA_x^\prime}{2\hbar}a^2+\mathcal{O}(a^3) }$

where for brevity with denoted: $\theta=\theta^x_{m,n},~~A_x=\theta^\prime=\partial_a\theta^x_{m,n}\big\vert_{a=0}$ and $A_x^\prime=\theta^{\prime\prime}=\partial^2_a\theta^x_{m,n}\big\vert_{a=0}$. Substituting these expansions to relevant part of the Hamiltonian yields

${\displaystyle \begin{align} &e^{i\theta}\vert m+a\rangle\langle m\vert +e^{-i\theta}\vert m\rangle\langle m+a\vert=\\ &\bigg(1+\frac{iA_x}{\hbar}a-\frac{A_x^2}{2\hbar^2}a^2 +\frac{iA^\prime_x}{2\hbar}a^2+\mathcal{O}(a^3)\bigg)\bigg(1-\frac{i\mathbf{p}_x}{\hbar}a-\frac{\mathbf{p}_x^2}{2\hbar^2}a^2 +\mathcal{O}(a^3)\bigg)\vert m\rangle\langle m\vert+\text{h.c}= \\&\bigg( 2-\frac{\mathbf{p}^2_x}{\hbar^2}a^2+\frac{e\lbrace \mathbf{p}_x,A_x \rbrace}{\hbar^2}a^2-\frac{e^2A_x^2}{\hbar^2}a^2+\mathcal{O}(a^3)\bigg) \vert m\rangle\langle m\vert\\ &\approx\bigg(-\frac{a^2}{\hbar^2}\big(\mathbf{p}_x-eA_x\big)^2+2+\mathcal{O}(a^3)\bigg) \vert m\rangle\langle m\vert \end{align} }$

Generalizing the last result to the 2D case, the we arrive to Hofstader Hamiltonian at the continuum limit:

$H_0=\frac{1}{2m}\big(\mathbf{p}-e\mathbf{A}\big)^2+\tilde{\epsilon_0}$

where the effective mass is $m=\hbar^2/2ta^2$ and $\tilde{\epsilon}_0=\epsilon_0+4$.

The semi-classical approach

Here we show that the Peierls phase originates from the propagator of an electron in a magnetic field due to the dynamical term $q\mathbf{v}\cdot\mathbf{A}$ appearing in the Lagrangian. In the path integral formalism, which generalizes the action principle of classical mechanics, the transition amplitude from site $j$ at time $t_j$ to site $i$ at time $t_i$ is given by

${\displaystyle \langle\mathbf{r}_i,t_i|\mathbf{r}_j,t_j\rangle = \int_{\mathbf{r}(t_i)}^{\mathbf{r}(t_j)} \mathcal{D}[\mathbf{r}(t)]e^{\frac{\rm i}{\hbar}\mathcal{S}(\mathbf{r})},}$

where the integration operator, $\int_{\mathbf{r}(t_i)}^{\mathbf{r}(t_j)} \mathcal{D}[\mathbf{r}(t)]$ denotes the sum over all possible paths from $\mathbf{r}(t_i)$ to $\mathbf{r}(t_j)$ and $\mathcal{S}[x]=\int_0^T L[x(t)] \mathrm{d}t$ is the classical action. The Lagrangian of the system can be written as

${\displaystyle L = L^{(0)}+q\mathbf{v}\cdot\mathbf{A}, }$

where $L^{(0)}$ is the Lagrangian in the absence of a magnetic field. The corresponding action reads

${\displaystyle S(\mathbf{r}_i,\mathbf{r}_j)=S^{(0)}(\mathbf{r}_i,\mathbf{r}_j)+q\int_{t_i}^{t_j}dt\left(\frac{\text{d}\mathbf{r}}{\text{d}t}\right)\cdot\mathbf{A}= S^{(0)}(\mathbf{r}_i,\mathbf{r}_j)+q\int_{\mathbf{r}_i}^{\mathbf{r}_j}\mathbf{A}\cdot\text{d}\mathbf{r} }$

Now, Assuming that all possible paths but the one joining $i$ and $j$ are neglectable we get

${\displaystyle \langle\mathbf{r}_i,t_i|\mathbf{r}_j,t_j\rangle=\int_{\mathbf{r}(t_i)}^{\mathbf{r}(t_j)} \mathcal{D}[\mathbf{r}(t)]e^{\frac{\rm i}{\hbar}\mathcal{S}^{(0)}(\mathbf{r})} e^{\frac{iq}{\hbar}\int_{\mathbf{r}_i}^{\mathbf{r}_j}\mathbf{A}\cdot\text{d}\mathbf{r}}}$

Hence, the transition amplitude of an electron subject to a magnetic field is the one in the absence of a magnetic field times a phase.

A rigorous derivation

The Hamiltonian is given by

${\displaystyle H=\frac{\mathbf{p}^2}{2m}+U\left(\mathbf{r}\right), }$ where $U\left(\mathbf{r}\right)$

is the potential landscape due to the crystal lattice. The Bloch theorem asserts that the solution to the problem

$H\Psi_{\mathbf{k}}=E\left(\mathbf{k}\right)\Psi_{\mathbf{k}},$

is to be sought in the Bloch sum form

${\displaystyle \Psi_{\mathbf{k}}=\frac{1}{\sqrt{N}}\sum_{\mathbf{R}}e^{i\mathbf{k}\cdot\mathbf{R}}\phi_\mathbf{R}\left(\mathbf{r}\right), }$

where $N$ is the number of unit cells, and $\phi$ are known as Wannier states. The corresponding eigenvalues $E\left(\mathbf{k}\right)$, which form bands depending on the crystal momentum $\mathbf{k}$, are obtained by calculating the matrix element $\langle\Psi_{\mathbf{k}}|H|\Psi_{\mathbf{k}}\rangle$

${\displaystyle \frac{1}{N}\sum_{\mathbf{R}\mathbf{R}^{\prime}}e^{i\mathbf{k}\left(\mathbf{R}^{\prime}-\mathbf{R}\right)} \int d\mathbf{r}\phi^*_\mathbf{R}\left(\mathbf{r}\right)H\phi_\mathbf{R}^{\prime}\left(\mathbf{r}\right) }$

and ultimately depend on material-related hopping integrals

${\displaystyle t_{12}=-\int d\mathbf{r}\phi^*_{\mathbf{R}_1}\left(\mathbf{r}\right)H\phi_{\mathbf{R}_2}\left(\mathbf{r}\right).}$

In the presence of the magnetic field the Hamiltonian changes to

${\displaystyle H'(t)=\frac{\left(\mathbf{p}-q\mathbf{A}(t)\right)^2}{2m}+U\left(\mathbf{r}\right), }$

where $q$ is the charge of the particle. To amend this consider changing the Wannier states to

${\displaystyle \begin{align} |\phi_\mathbf{R}'(\mathbf{r})\rangle = e^{i \frac{q}{\hbar} \int_\mathbf{R}^\mathbf{r} \mathbf{A}(\mathbf{r}',t) \cdot dr'} |\phi_\mathbf{R}(\mathbf{r})\rangle, \end{align} }$

where $|\phi_\mathbf{R}> \equiv |\phi_\mathbf{R}'(\mathbf{A}\to 0)\rangle$. This makes the new Bloch wave functions

${\displaystyle |\Psi_\mathbf{k}'\rangle = \frac{1}{\sqrt{N}} \sum_{\mathbf{R}} e^{i \mathbf{k}\cdot\mathbf{R}} |\phi_\mathbf{R}'(\mathbf{r})\rangle, }$

into eigenstates of the full Hamiltonian at time $t$, with the same energy as before. To see this we first use $\mathbf{p} = i \hbar \nabla$ to write

${\displaystyle \begin{align} H'(t) |{\phi_\mathbf{R}'(\mathbf{r})}\rangle &= \left[ \frac{(\mathbf{p} - q\mathbf{A}(\mathbf{r},t))^2}{2m} + U(\mathbf{r}) \right] e^{i\frac{q}{\hbar} \int_\mathbf{R}^\mathbf{r} \mathbf{A}(\mathbf{r}',t) \cdot d\mathbf{r}'} |\phi_\mathbf{R}(\mathbf{r})\rangle \\ & = e^{i\frac{q}{\hbar} \int_\mathbf{R}^\mathbf{r} A(\mathbf{r}',t) \cdot d\mathbf{r}'} \left[\frac{(\mathbf{p} - q\mathbf{A}(\mathbf{r},t) + q \mathbf{A}(\mathbf{r},t))^2}{2m} + U(\mathbf{r}) \right] |\phi_\mathbf{R}\rangle \\ & = e^{i\frac{q}{\hbar} \int_\mathbf{R}^\mathbf{r} A(\mathbf{r}',t) \cdot d\mathbf{r}'} H |\phi_\mathbf{R}\rangle. \end{align}}$

Then when we compute the hopping integral in quasi-equilibrium (assuming that the vector potential changes slowly)

${\displaystyle \begin{align} t_{\mathbf{R}\mathbf{R}'}(t)&=-\int d\mathbf{r} \langle\phi_\mathbf{R}'(\mathbf{r})|H'(t)|\phi_{\mathbf{R}'}'(\mathbf{r})\rangle \\ &= - \int d\mathbf{r} \langle \phi_\mathbf{R}(\mathbf{r})|e^{i\frac{q}{\hbar} \left[-\int_{\mathbf{R}'}^\mathbf{r} \mathbf{A}(\mathbf{r}',t)\cdot d\mathbf{r}'+\int_\mathbf{R}^\mathbf{r} \mathbf{A}(\mathbf{r}',t)\cdot d\mathbf{r}'\right] } H | \phi_{\mathbf{R}'} \rangle \\ & = - e^{i\frac{q}{\hbar}\int_{\mathbf{R}}^{\mathbf{R}'} \mathbf{A}(\mathbf{r}',t)\cdot d\mathbf{r}' } \int d\mathbf{r} \langle \phi_\mathbf{R}(\mathbf{r})| e^{i\frac{q}{\hbar}\Phi_{\mathbf{R},\mathbf{r},\mathbf{R}'}} H | \phi_{\mathbf{R}'} \rangle, \end{align} }$

where we have defined $\Phi_{\mathbf{R},\mathbf{r},\mathbf{R}'} = \int_{\mathbf{R}\to \mathbf{r} \to \mathbf{R}' \to \mathbf{R}}\mathbf{A}(\mathbf{r}',t)\cdot d\mathbf{r}'$, the flux through the triangle made by the three position arguments. Since we assume $\mathbf{A}(\mathbf{r},t)$ is approximately uniform at the lattice scale^[4] - the scale at which the Wannier states are localized to the positions $\mathbf{R}$ - we can approximate $\Phi_{\mathbf{R},\mathbf{r},\mathbf{R}'} \approx 0$, yielding the desired result, ${\displaystyle t_{\mathbf{R}\mathbf{R}'}(t) \approx t_{\mathbf{R}\mathbf{R}'} e^{i\frac{q}{\hbar}\int_{\mathbf{R}}^{\mathbf{R}'} \mathbf{A}(\mathbf{r}',t)\cdot d\mathbf{r}' }. }$ Therefore the matrix elements are the same as in the case without magnetic field, apart from the phase factor picked up, which is denoted the Peierls phase. This is tremendously convenient, since then we get to use the same material parameters regardless of the magnetic field value, and the corresponding phase is computationally trivial to take into account. For electrons it amounts to replacing the hopping term $t_{ij}$ with $t_{ij}e^{i\frac{e}{\hbar}\int_i^j\mathbf{A}\cdot d\mathbf{l}}$ ^{[5] [6] [7]} .^[4]

References

1. Peierls, R (1933). On the theory of diamagnetism of conduction electrons. Z. Phys 80 (World Scientific). pp. 763–791.
2. Hofstadter, Douglas R. (Sep 1976). Energy levels and wave functions of Bloch electrons in rational and irrational magnetic fields. Phys. Rev. B 14 (American Physical Society). pp. 2239–2249. doi:10.1103/PhysRevB.14.2239.
3. Feynman Richard P Sands Matthew L Leighton Robert B; Richard Phillips Feynman; Robert B. Leighton; Matthew Linzee Sands (25 November 2013). The Feynman Lectures on Physics, Desktop Edition Volume III: The New Millennium Edition. Basic Books. pp. 9–. ISBN 978-0-465-07997-1.
4. Luttinger, J. M. (Nov 1951). The Effect of a Magnetic Field on Electrons in a Periodic Potential. Phys. Rev. 84 (American Physical Society). pp. 814–817. doi:10.1103/PhysRev.84.814.
5. Kohn, Walter (Sep 1959). Theory of Bloch Electrons in a Magnetic Field: The Effective Hamiltonian. Phys. Rev. 115 (American Physical Society). pp. 1460–1478. doi:10.1103/PhysRev.115.1460.
6. Blount, E. I. (Jun 1962). Bloch Electrons in a Magnetic Field. Phys. Rev. 126 (American Physical Society). pp. 1636–1653. doi:10.1103/PhysRev.126.1636.
7. Wannier, Gregory H. (Oct 1962). Dynamics of Band Electrons in Electric and Magnetic Fields. Rev. Mod. Phys. 34 (American Physical Society). pp. 645–655. doi:10.1103/RevModPhys.34.645.