What is Wannier? --the wisdom of simplification and localization

Mon, 10 Aug 2026 00:00:00 GMT

This article is a simple introduction to the Wannier function. It won’t go into more technical details, but will give a general idea of what it is and what it does in the practical calculation with my own understanding and thinking framework. It’s not a formal article and may contain some mistakes, feel free to contact me if you have any questions or suggestions.

Where does it come from?#

The modern condensed matter physics aims to figure out the electron structures and behaviours in the crystal-like, i.e. the periodic potential field. As a fundamental theory, Bloch’s theorem tells the most important property of the wave function while it is noticed that the crystal momentum kkk is a good quantum number in such periodic potential. [1]

ψnk(r+R)=eik⋅Rψnk(r).\begin{align} \psi_{n\mathbf k}(\mathbf r+\mathbf R) &= e^{i\mathbf k\cdot\mathbf R} \psi_{n\mathbf k}(\mathbf r). \end{align}ψnk(r+R)=eik⋅Rψnk(r).

However, Bloch’s picture describe the electron as a “running wave” in the infinite potential field, which is not naturally consistent with the “atomic” picture. Early quantum mechanics provided a clear orbital picture for isolated atoms, most notably through the exact solution of the hydrogen atom and approximate treatments of many-electron atoms. It is simpler and more intuitive to treat the electron states in the periodic potential as the linear combination of the localized atomic-orbital-like real space functions.

Gregory Wannier gave his answer in 1937 when studying the electron-hole excitation in the crystal. He constructed the Wannier Functions (WFs) to bridge the above gap. [2]

Wannier functions are real-space representatives of a chosen Bloch-band subspace. For a finite crystal containing NNN unit cells with periodic boundary conditions, the Wannier states associated with an isolated band are defined by

∣wnR⟩=1N∑ke−ik⋅R∣ψnk⟩.\begin{align} \left|w_{n\mathbf R}\right\rangle &= \frac{1}{\sqrt{N}} \sum_{\mathbf k} e^{-i\mathbf k\cdot\mathbf R} \left|\psi_{n\mathbf k}\right\rangle. \end{align}∣wnR⟩=N1k∑e−ik⋅R∣ψnk⟩.

For a composite subspace containing JJJ bands, the Bloch states at each k\mathbf kk can also mix with each other:

∣wnR⟩=1N∑ke−ik⋅R∑m=1J∣ψmk⟩Umn(k).\begin{align} |w_{n\mathbf R}\rangle &= \frac{1}{\sqrt N} \sum_{\mathbf k} e^{-i\mathbf k\cdot\mathbf R} \sum_{m=1}^{J} |\psi_{m\mathbf k}\rangle U_{mn}(\mathbf k). \end{align}∣wnR⟩=N1k∑e−ik⋅Rm=1∑J∣ψmk⟩Umn(k).

Here U(k)∈U(J)U(\mathbf k)\in U(J)U(k)∈U(J) is a k\mathbf kk-dependent unitary matrix. It represents the gauge freedom inside the selected band subspace. Different choices of U(k)U(\mathbf k)U(k) keep the same subspace, but generally give different Wannier functions and different degrees of localization.

Their position-space representations are complex-valued functions:

wnR(r)=⟨r|wnR⟩,\begin{align} w_{n\mathbf R}(\mathbf r) &= \left\langle\mathbf r\middle|w_{n\mathbf R}\right\rangle, \end{align}wnR(r)=⟨r∣wnR⟩,

Wannier functions have the following important properties.

Lattice-translation covariance

Wannier functions in different unit cells are related by lattice translations:
wnR(r)=wn0(r−R).\begin{align} w_{n\mathbf R}(\mathbf r) &= w_{n\mathbf 0}(\mathbf r-\mathbf R). \end{align}wnR(r)=wn0(r−R).
Orthonormality
⟨wnR|wmR′⟩=δnmδRR′\begin{align} \left\langle w_{n\mathbf R} \middle| w_{m\mathbf R'} \right\rangle &= \delta_{nm}\delta_{\mathbf R\mathbf R'} \end{align}⟨wnR∣wmR′⟩=δnmδRR′
Completeness within the selected band subspace

The Wannier states form an orthonormal basis for the range of a band projector P^\hat PP^:
HP:=Ran⁡P^=span⁡‾{∣wnR⟩}n,R\begin{align} \mathcal H_P &:= \operatorname{Ran}\hat P = \overline{\operatorname{span}} \left\{ \left|w_{n\mathbf R}\right\rangle \right\}_{n,\mathbf R} \end{align}HP:=RanP^=span{∣wnR⟩}n,R
The projector satisfies
P^2=P^, P^†=P^P^=∑n,R∣wnR⟩⟨wnR∣\begin{aligned} &\hat P^2= \hat P, \ \hat P^\dagger = \hat P \\ &\hat P = \sum_{n,\mathbf R} \left|w_{n\mathbf R}\right\rangle \left\langle w_{n\mathbf R}\right| \end{aligned}P^2=P^, P^†=P^P^=n,R∑∣wnR⟩⟨wnR∣
Real-space localization

When a sufficiently smooth gauge exists, the Wannier functions can be chosen to decay rapidly away from the centers. For exponentially localized Wannier functions,
∣wnR(r)∣≤Cne−αn∣r−R∣,Cn,αn>0.\begin{align} \left| w_{n\mathbf R}(\mathbf r) \right| &\leq C_n e^{-\alpha_n \left|\mathbf r-\mathbf R\right|}, \qquad C_n,\alpha_n>0. \end{align}∣wnR(r)∣≤Cne−αn∣r−R∣,Cn,αn>0.

Therefore, Wannier functions do not satisfy the Bloch condition. For an isolated band, the Bloch state is obtained as a momentum-labelled superposition of Wannier functions distributed over all lattice cells:

∣ψnk⟩=1N∑Reik⋅R∣wnR⟩.\begin{align} \left|\psi_{n\mathbf k}\right\rangle &= \frac{1}{\sqrt N} \sum_{\mathbf R} e^{i\mathbf k\cdot\mathbf R} \left|w_{n\mathbf R}\right\rangle. \end{align}∣ψnk⟩=N1R∑eik⋅R∣wnR⟩.

For composite bands, the inverse transformation instead reconstructs a Bloch-like state in the Wannier gauge,

∣ψ~nk⟩=∑m=1J∣ψmk⟩Umn(k)=1N∑Reik⋅R∣wnR⟩.\begin{align} \left|\widetilde\psi_{n\mathbf k}\right\rangle &= \sum_{m=1}^{J} \left|\psi_{m\mathbf k}\right\rangle U_{mn}(\mathbf k) \notag \\ &= \frac{1}{\sqrt N} \sum_{\mathbf R} e^{i\mathbf k\cdot\mathbf R} \left|w_{n\mathbf R}\right\rangle. \end{align}ψnk⟩=m=1∑J∣ψmk⟩Umn(k)=N1R∑eik⋅R∣wnR⟩.

The two representations, the Bloch ∣ψnk⟩\ket{\psi_{n\mathbf k}}∣ψnk⟩ and the Wannier ∣wnR⟩\ket{w_{n\mathbf R}}∣wnR⟩ both describe the correct band dispersion. Bloch states are extended throughout the crystal, whereas Wannier states sacrifice a definite momentum label in exchange for real-space localization.

What does it do?#

In the condensed matter calculation, Wannierization is most commonly used as a post-processing step following a first-principles electronic structure calculation, mostly Density Functional Theory (DFT), even though it is not the initial purpose of Wannier’s original work.

In 1997, Marzari and Vanderbilt introduced the maximally localized Wannier functions (MLWFs) for composite bands [3]. In 2001, Souza, Marzari, and Vanderbilt extended this method to entangled bands through a disentanglement procedure [4]. In 2008, the Wannier90 package was published [5].

Wannier90 does not determine the electronic structure by itself. Instead, it takes the eigenvalues, Bloch-state overlaps, and trial-orbital projections supplied by a first-principles calculation and constructs a localized representation of a selected band subspace. Within a fixed isolated subspace, Wannierization is a unitary change of basis. Approximations enter when an effective subspace is disentangled and when long-ranged real-space matrix entries are truncated.

In a first-principles workflow, DFT supplies the Kohn-Sham Hamiltonian, eigenvalues, and Bloch eigenstates:

H^KS[ρ]∣ψnk⟩=εnk∣ψnk⟩.\begin{align} \hat H_{\mathrm{KS}}[\rho] \left|\psi_{n\mathbf k}\right\rangle &= \varepsilon_{n\mathbf k} \left|\psi_{n\mathbf k}\right\rangle. \end{align}H^KS[ρ]∣ψnk⟩=εnk∣ψnk⟩.

Wannierization gives the Real Space Hamiltonian under the Real Space Wannier basis

Hmn(R)=⟨wm0∣H^∣wnR⟩\begin{align} H_{mn}(\mathbf R) = \bra{w_{m\mathbf 0}}\hat{H}\ket{w_{n\mathbf R}} \end{align}Hmn(R)=⟨wm0∣H^∣wnR⟩

And the Fourier transformation to momentum reciprocal space

Hmn(k)=∑Reik⋅RHmn(R).\begin{align} H_{mn}(\mathbf k) &= \sum_{\mathbf R} e^{i\mathbf k\cdot\mathbf R} H_{mn}(\mathbf R). \end{align}Hmn(k)=R∑eik⋅RHmn(R).

So why do we wannierize it instead of using the original DFT result directly?

First-principles DFT calculations provide an accurate description of the electronic structure. In principle, most physical quantities can be evaluated directly from the Kohn-Sham eigenstates. In practice, the difficulty lies in repeating such calculations over the extremely dense momentum meshes required for Brillouin-zone integration. Such large plane-wave diagonalizations on the mesh is too expensive to be practical, especially for large systems.

Wannierization replaces large-scale calculation with a compact interpolation problem. Dense-mesh calculations require only the diagonalization of a much smaller Wannier Hamiltonian. Topological invariants, Berry curvature, transport coefficients, and dynamical response functions can therefore be evaluated efficiently without solving the original large plane-wave problem.

For example, the primitive bcc Fe unit cell contains only 1 magnetic atom, but the DFT calculation requires more than 600 plane waves to converge the Kohn-Sham eigenstates. In contrast, a Wannier Hamiltonian with 18 basis functions per unit cell is sufficient to reproduce the correct band structure over a wide energy range centered on the Fermi level. The Wannier Hamiltonian is therefore more than an order of magnitude smaller than the original DFT Hamiltonian, and its diagonalization is correspondingly faster.

The bcc Fe band structure, The grey lines represent the vasp results, and the red lines represent the Wannier90 results

1-dimensional ssh-like example#

Suppose we have a 1-dimensional SSH-like model with two identical atoms in a unit cell.

periodic potential

lattice constant aaa
relative position ddd

The potential is given by

V(x)=−V0∑R∈Λ[g(x−R)+g(x−R−d)]\begin{align} V(x) &= -V_0 \sum_{R\in \Lambda} [g(x-R) + g(x-R-d)] \end{align}V(x)=−V0R∈Λ∑[g(x−R)+g(x−R−d)]

where g(x)g(x)g(x) is the Gaussian function defined as

g(x)=e−x22σ2.\begin{align} g(x) &= e^{-\frac{x^2}{2\sigma^2}}. \end{align}g(x)=e−2σ2x2.

and the lattice set Λ\LambdaΛ is defined as

Λ={ma∣m∈Z}.\begin{align} \Lambda &= \{ma| m\in \mathbb Z\}. \end{align}Λ={ma∣m∈Z}.

Plane wave solution#

The Hamiltonian is given by

H^=−ℏ22md2dx2+V(x)\begin{align} \hat H &= -\frac{\hbar^2}{2m}\frac{d^2}{dx^2} + V(x) \end{align}H^=−2mℏ2dx2d2+V(x)

from the Bloch theorem, the Bloch wave function is the eigenfunction of the Hamiltonian.

ψnk(x)=eikxunk(x)\begin{align} \psi_{n k}(x) = e^{ikx}u_{nk}(x) \end{align}ψnk(x)=eikxunk(x)

expand the periodic function unk(x)u_{nk}(x)unk(x)

ψnk(x)=∑GCnG(k)ei(k+G)x.\begin{align} \psi_{nk}(x) &= \sum_G C_{nG}(k)e^{i(k+G)x}. \end{align}ψnk(x)=G∑CnG(k)ei(k+G)x.

We can define the normalized plane wave basis in the real space as

⟨x∣k+G⟩=1Lei(k+G)x.\begin{align} \braket{x|k+G} &= \frac{1}{\sqrt L}e^{i(k+G)x}. \end{align}⟨x∣k+G⟩=L1ei(k+G)x.

Then we have the Bloch states in the plane wave basis as

∣ψnk⟩=∑GCnG(k)∣k+G⟩.\begin{align} \ket{\psi_{nk}} &= \sum_G C_{nG}(k)\ket{k+G}. \end{align}∣ψnk⟩=G∑CnG(k)∣k+G⟩.

Then the eigenvalue problem can be written as

∑G′⟨k+G∣H^∣k+G′⟩CnG′(k)=EnkCnG(k).\begin{align} \sum_{G'}\bra{k+G}\hat H\ket{k+G'}C_{nG'}(k) &= E_{nk}C_{nG}(k). \end{align}G′∑⟨k+G∣H^∣k+G′⟩CnG′(k)=EnkCnG(k).

Kinetic energy#

⟨k+G∣T^∣k+G′⟩=ℏ22m(k+G)2δGG′\begin{align} \bra{k+G}\hat T\ket{k+G'} &= \frac{\hbar^2}{2m}(k+G)^2\delta_{GG'} \end{align}⟨k+G∣T^∣k+G′⟩=2mℏ2(k+G)2δGG′

which is the diagonal term

Potential energy#

⟨k+G∣V^∣k+G′⟩=VG−G′\begin{align} \bra{k+G}\hat V\ket{k+G'} &= V_{G-G'} \end{align}⟨k+G∣V^∣k+G′⟩=VG−G′

In order to find the off-diagonal term, we can use the Fourier transformation to get the potential in the reciprocal space

VG=1a∫0aV(x)e−iGxdx\begin{align} V_G &= \frac{1}{a}\int_0^a V(x)e^{-iGx} dx \end{align} VG=a1∫0aV(x)e−iGxdxVG=−V0a∑R∈Λ∫0a[g(x−R)+g(x−R−d)]e−iGxdx\begin{aligned} V_G= -\frac{V_0}{a}\sum_{R\in \Lambda} \int_0^a [g(x-R) + g(x-R-d)] e^{-iGx} dx \end{aligned}VG=−aV0R∈Λ∑∫0a[g(x−R)+g(x−R−d)]e−iGxdx

here we only show the contribution from the first term, the second term is similar

VG(1)=−V0a∫0a∑R∈Λg(x−R)e−iGxdx=−V0a∑m∈Z∫0ae−(x−ma)22σ2e−iGxdx=−V0a∑m∈Z∫−maa−mae−x22σ2e−iG(x+ma)dx\begin{aligned} V_G^{(1)} &= -\frac{V_0}{a}\int_0^a \sum_{R\in \Lambda} g(x-R) e^{-iGx} dx \\ &= -\frac{V_0}{a} \sum_{m \in \mathbb{Z}} \int_{0}^{a} e^{-\frac{(x-ma)^2}{2\sigma^2}} e^{-iGx} dx \\ &= -\frac{V_0}{a} \sum_{m \in \mathbb{Z}} \int_{-ma}^{a-ma} e^{-\frac{x^2}{2\sigma^2}} e^{-iG(x+m a)} dx \\ \end{aligned}VG(1)=−aV0∫0aR∈Λ∑g(x−R)e−iGxdx=−aV0m∈Z∑∫0ae−2σ2(x−ma)2e−iGxdx=−aV0m∈Z∑∫−maa−mae−2σ2x2e−iG(x+ma)dx

mmm enumerates all the integers

e−iGma=e−i2πajma=1e^{-iGma} = e^{-i\frac{2\pi}{a}jm a} = 1e−iGma=e−ia2πjma=1VG(1)=−V0a∫−∞∞e−x22σ2e−iGxdx=−V0a2πσe−G2σ22V_G^{(1)} = -\frac{V_0}{a} \int_{-\infty}^{\infty} e^{-\frac{x^2}{2\sigma^2}} e^{-iGx} dx = -\frac{V_0}{a} \sqrt{2\pi}\sigma e^{-\frac{G^2\sigma^2}{2}}VG(1)=−aV0∫−∞∞e−2σ2x2e−iGxdx=−aV02πσe−2G2σ2

The second term is similar, we can get the final result

VG=−V0a2πσe−G2σ22(1+e−iGd)\begin{align} V_G &= -\frac{V_0}{a} \sqrt{2\pi}\sigma e^{-\frac{G^2\sigma^2}{2}}(1+e^{-iGd}) \end{align}VG=−aV02πσe−2G2σ2(1+e−iGd)

so the potential entry is

VG−G′=−V0a2πσe−(G−G′)2σ22(1+e−i(G−G′)d)\begin{align} V_{G-G'} &= -\frac{V_0}{a} \sqrt{2\pi}\sigma e^{-\frac{(G-G')^2\sigma^2}{2}}(1+e^{-i(G-G')d}) \end{align}VG−G′=−aV02πσe−2(G−G′)2σ2(1+e−i(G−G′)d)

with the Hamiltonian entry

⟨k+G∣H^∣k+G′⟩=−V0a2πσe−(G−G′)2σ22(1+e−i(G−G′)d)+ℏ22m(k+G)2δGG′\begin{align} \bra{k+G}\hat H\ket{k+G'} &= -\frac{V_0}{a} \sqrt{2\pi}\sigma e^{-\frac{(G-G')^2\sigma^2}{2}}(1+e^{-i(G-G')d}) + \frac{\hbar^2}{2m}(k+G)^2\delta_{GG'} \end{align}⟨k+G∣H^∣k+G′⟩=−aV02πσe−2(G−G′)2σ2(1+e−i(G−G′)d)+2mℏ2(k+G)2δGG′

we can construct the Hamiltonian matrix and diagonalize it at any given kkk point to get the band structure. We take total 111 dimension with G=−552πa,−542πa,⋯ ,552πaG = -55\frac{2\pi}{a}, -54\frac{2\pi}{a}, \cdots, 55\frac{2\pi}{a}G=−55a2π,−54a2π,⋯,55a2π, the proper parameters are selected. We get the lowest band structure as shown in the figure below.

band structure

Tight-binding approximation#

In practical calculation, we only care about the specific bands of the system, which is usually near the Fermi surface, or with nontrivial topological structures by selecting an energy window to construct a subspace. The lowest two plane-wave bands are separated from the higher-energy bands by a finite gap and therefore form a well-defined two-dimensional invariant subspace. We would like to describe this subspace using two localized orbitals per unit cell.

Suppose we have 2 orbitals in the 2 positions of the unit cell as ∣A,R⟩\ket{A, R}∣A,R⟩ and ∣B,R⟩\ket{B, R}∣B,R⟩, we can write Hopping term

Onsite ⟨A,R∣H^∣A,R⟩=⟨B,R∣H^∣B,R⟩=ϵ0\begin{align} \bra{A, R}\hat H\ket{A, R} &= \bra{B, R}\hat H\ket{B, R} = \epsilon_0 \end{align}⟨A,R∣H^∣A,R⟩=⟨B,R∣H^∣B,R⟩=ϵ0
Inter-cell same site ⟨A,R∣H^∣A,R+a⟩=⟨B,R∣H^∣B,R+a⟩=t0\begin{align} \bra{A, R}\hat H\ket{A, R+a} &= \bra{B, R}\hat H\ket{B, R+a} = t_0 \end{align}⟨A,R∣H^∣A,R+a⟩=⟨B,R∣H^∣B,R+a⟩=t0
Intra-cell ⟨A,R∣H^∣B,R⟩=t1\begin{align} \bra{A, R}\hat H\ket{B, R} &= t_1 \end{align}⟨A,R∣H^∣B,R⟩=t1
Inter-cell different site ⟨A,R+a∣H^∣B,R⟩=t2\begin{align} \bra{A, R+a}\hat H\ket{B, R} &= t_2 \end{align}⟨A,R+a∣H^∣B,R⟩=t2

Second quantization form

H^ext=∑Rϵ0(c^A,R†c^A,R+c^B,R†c^B,R)+∑R[t0(c^A,R+a†c^A,R+c^B,R+a†c^B,R)+h.c.]+∑R[t1c^A,R†c^B,R+t2c^A,R+a†c^B,R+h.c.].\begin{align} \hat H_{\mathrm{ext}} ={}& \sum_R \epsilon_0 \left( \hat c_{A,R}^\dagger\hat c_{A,R} + \hat c_{B,R}^\dagger\hat c_{B,R} \right) + \sum_R \left[ t_0 \left( \hat c_{A,R+a}^\dagger\hat c_{A,R} + \hat c_{B,R+a}^\dagger\hat c_{B,R} \right) +\mathrm{h.c.} \right] \notag \\ &+ \sum_R \left[ t_1\hat c_{A,R}^\dagger\hat c_{B,R} + t_2\hat c_{A,R+a}^\dagger\hat c_{B,R} + \mathrm{h.c.} \right]. \end{align}H^ext=R∑ϵ0(c^A,R†c^A,R+c^B,R†c^B,R)+R∑[t0(c^A,R+a†c^A,R+c^B,R+a†c^B,R)+h.c.]+R∑[t1c^A,R†c^B,R+t2c^A,R+a†c^B,R+h.c.].

Fourier transforming the annihilation and creation operators

c^A,R=1N∑keikRc^A,kc^B,R=1N∑keikRc^B,k\begin{align} \hat c_{A,R} &= \frac{1}{\sqrt N} \sum_k e^{ikR} \hat c_{A,k} \notag \\ \hat c_{B,R} &= \frac{1}{\sqrt N} \sum_k e^{ikR} \hat c_{B,k} \end{align}c^A,Rc^B,R=N1k∑eikRc^A,k=N1k∑eikRc^B,k

We can write the Hamiltonian in the momentum space by

H^=∑kΨk†H(k)Ψk\begin{align} \hat H = \sum_k \Psi_k^\dagger H(k) \Psi_k \end{align}H^=k∑Ψk†H(k)Ψk

where Ψk=[c^A,k,c^B,k]T\Psi_k = [\hat c_{A,k}, \hat c_{B,k}]^TΨk=[c^A,k,c^B,k]T and the Hamiltonian matrix is

H(k)=[ϵ0+2t0cos⁡(ka)t1+t2e−ikat1+t2eikaϵ0+2t0cos⁡(ka)]\begin{align} H(k) &= \begin{bmatrix} \epsilon_0 + 2t_0\cos(ka) & t_1 + t_2 e^{-ika} \\ t_1 + t_2 e^{ika} & \epsilon_0 + 2t_0\cos(ka) \end{bmatrix} \end{align}H(k)=[ϵ0+2t0cos(ka)t1+t2eikat1+t2e−ikaϵ0+2t0cos(ka)]

(ϵ0,t0,t1,t2∈R\epsilon_0, t_0, t_1, t_2 \in \mathbb{R}ϵ0,t0,t1,t2∈R)

With the proper parameters, we can get the band structure as shown below.

TB band structure

Where Wannier Enters#

The two calculations above reveal the trade-off between the plane-wave and tight-binding descriptions. The plane-wave method solves the continuum Hamiltonian in a large and delocalized basis. However, the resulting Hamiltonian is large and its real-space interpretation is not immediately transparent.

The extended SSH model takes the opposite approach. It assumes two localized orbitals, ∣A,R⟩\ket{A,R}∣A,R⟩ and ∣B,R⟩\ket{B,R}∣B,R⟩, and retains only a small number of hopping parameters. The resulting 2×22\times22×2 Hamiltonian is simple and physically intuitive. However, the model itself does not tell us what the two orbital wave functions actually are, nor does it determine the values of ϵ0,t0,t1,\epsilon_0,t_0,t_1,ϵ0,t0,t1, and t2t_2t2. In the example above, these parameters were fitted manually to the plane-wave band structure.

Wannierization provides the missing connection between these two descriptions. Starting from the isolated two-band subspace obtained in the plane-wave calculation, it constructs two localized Wannier functions per unit cell. In the MLWF method, the gauge freedom of the Bloch states is chosen by minimizing the real-space spread of these functions. The states ∣A,R⟩\ket{A,R}∣A,R⟩ and ∣B,R⟩\ket{B,R}∣B,R⟩ in the tight-binding model can therefore be replaced by explicitly constructed Wannier functions rather than assumed atomic-like orbitals.

Once the Wannier basis has been constructed, the original Hamiltonian is represented by the real-space matrix elements as shown in Eq. (12).

The onsite energies and hopping parameters are then identified from these matrix elements. For example,

ϵ0∼H11(0),t1∼H12(0),t0∼H11(a),t2∼H12(−a).\begin{align} \epsilon_0 &\sim H_{11}(0), & t_1 &\sim H_{12}(0), \notag \\ t_0 &\sim H_{11}(a), & t_2 &\sim H_{12}(-a). \end{align}ϵ0t0∼H11(0),∼H11(a),t1t2∼H12(0),∼H12(−a).

Unlike the SSH model above, the Wannier Hamiltonian also reveals longer-ranged hopping processes whenever their matrix elements are non-negligible. A practical tight-binding model is obtained by choosing a real-space cutoff and retaining the relevant matrix elements:

Hmn(k)=∑R∈cutoff seteik⋅RHmn(R).\begin{align} H_{mn}(k) &= \sum_{R \in \text{cutoff set}} e^{ik \cdot R} H_{mn}(R). \end{align}Hmn(k)=R∈cutoff set∑eik⋅RHmn(R).

Wannierization does more than provide another tight-binding model. It systematically derives the localized basis and its hopping parameters from the original plane-wave calculation, while allowing the accuracy and complexity of the effective model to be controlled through the real-space cutoff. With the real-space Wannier Hamiltonian, position operator, and other physical operators represented in the Wannier basis, we gain efficient access to these physical quantities. Wannierization does not change the physics; it changes the language in which the physics becomes transparent.

References:

1. F. Bloch, “Über die Quantenmechanik der Elektronen in Kristallgittern,” Z. Phys. 52, 555–600 (1929), DOI: 10.1007/BF01339455.

2. G. H. Wannier, “The Structure of Electronic Excitation Levels in Insulating Crystals,” Phys. Rev. 52, 191–197 (1937), DOI: 10.1103/PhysRev.52.191.

3. N. Marzari and D. Vanderbilt, “Maximally Localized Generalized Wannier Functions for Composite Energy Bands,” Phys. Rev. B 56, 12847–12865 (1997), DOI: 10.1103/PhysRevB.56.12847.

4. I. Souza, N. Marzari, and D. Vanderbilt, “Maximally Localized Wannier Functions for Entangled Energy Bands,” Phys. Rev. B 65, 035109 (2001), DOI: 10.1103/PhysRevB.65.035109.

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