Non-Hermitian XY Spin Chain: Theoretical Insights

• The non-Hermitian XY spin chain is a one-dimensional free-fermion model with complex anisotropies and imaginary field terms, leading to anti-linear symmetry and real spectra in unbroken phases.
• Analytical techniques such as the Jordan–Wigner transformation and complex Bogoliubov rotation yield explicit eigenvalue spectra, revealing exceptional points and non-Hermitian phase transitions.
• Topological invariants, entanglement measures, and observable prescriptions underscore unique quantum critical behaviors and phase transitions in these models.

Searching arXiv for recent and foundational papers on non-Hermitian XY spin chains to ground the article in the literature. arxiv_search(query="non-Hermitian XY spin chain", max_results=10, sort_by="relevance") The non-Hermitian XY spin chain is a class of one-dimensional spin-$\tfrac12$ lattice models in which the standard nearest-neighbor XY exchange is supplemented by complex anisotropies, complex transverse fields, staggered imaginary fields, or non-collinear non-Hermitian couplings. In the formulations studied on arXiv, these models remain closely tied to free-fermion structure after Jordan–Wigner transformation, yet they exhibit phenomena absent in Hermitian XY chains: anti-linear symmetry-unbroken regions with entirely real spectra, exceptional points (EPs) at which eigenvalues and eigenvectors coalesce, critical regions with pure-imaginary or complex gaps, biorthogonal versus standard expectation-value ambiguities, and non-Hermitian topological structures encoded in winding numbers, EP rings, and branch cuts of eigenstates (Li et al., 2022, Zhang et al., 2012, Luo et al., 5 Jun 2026).

1. Model family and anti-linear symmetry structure

A useful starting point is the nearest-neighbor non-Hermitian XY Hamiltonian in a transverse field,

$$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$

with real $\gamma$, $h$, and $\kappa$, from which two widely studied subclasses follow: a complex-field model $H_1$ with $h\to\lambda\in\mathbb C$, and an imaginary-anisotropy model $H_2$ with $\gamma\to i\kappa$ (Luo et al., 5 Jun 2026). Other concrete variants include an alternating imaginary transverse field

$$H= -\frac{J}{2}\sum_{\ell=1}^{N}\bigl[(1+\gamma)\sigma_{\ell}^{x}\sigma_{\ell+1}^{x} +(1-\gamma)\sigma_{\ell}^{y}\sigma_{\ell+1}^{y}\bigr] -h\sum_{\ell}\sigma_{\ell}^{z} +i\eta\sum_{\ell}(-1)^{\ell}\sigma_{\ell}^{z},$$

which is invariant under parity $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$0 and time reversal $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$1, hence $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$2-symmetric (Li et al., 2022), and the anisotropic chain with complex couplings

$$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$3

with $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$4, $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$5, which is not $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$6- but $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$7-symmetric, where $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$8 and $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$9 is complex conjugation (Zhang et al., 2012).

The anti-linear symmetry is model-dependent. In the $\gamma$0-symmetric chain of Zhang and Song, $\gamma$1 although $\gamma$2 and $\gamma$3 separately, and the spectrum is entirely real only when all eigenstates are $\gamma$4-symmetric (Zhang et al., 2012). In the alternating-field chain, the $\gamma$5-symmetric region is bounded by the onset of complex eigenvalues as the staggered imaginary field grows (Li et al., 2022). In the odd-length ring-frustrated chain, the non-Hermitian term is instead a symmetric non-collinear coupling,

$\gamma$6

which leads to a real spectrum only in specific parameter regions and is discussed in terms of $\gamma$7-symmetry breaking rather than $\gamma$8 or $\gamma$9 (Bi et al., 2020).

A recurring misconception is that commutation with an anti-linear operator suffices for spectral reality. In these XY chains, commutation with $h$0, $h$1, or $h$2 is only the algebraic prerequisite; the spectrum is real in the unbroken phase, where eigenvectors themselves are symmetry eigenstates, and complex eigenvalues appear once that symmetry is spontaneously broken (Zhang et al., 2012, Luo et al., 5 Jun 2026).

2. Exact solvability and fermionic reduction

Most non-Hermitian XY chains on arXiv retain a quadratic fermionic structure. The standard route is Jordan–Wigner transformation, followed by Fourier decomposition in parity sectors and a complex Bogoliubov transformation. In the $h$3-symmetric chain, the Jordan–Wigner map introduces spinless fermions $h$4, the even/odd fermion-number sectors are separated by projectors $h$5, and the Hamiltonian in each sector takes the diagonal form

$h$6

with single-particle dispersion

$h$7

The quasiparticles $h$8 are biorthogonal partners rather than Hermitian conjugates (Zhang et al., 2012).

For open chains with complex anisotropy $h$9, Li, Liu, and Batchelor formulate the problem through the $\kappa$0 quasi-Hamiltonian

$\kappa$1

where $\kappa$2 and $\kappa$3. The single-particle spectrum satisfies

$\kappa$4

and the boundary quantization can be written in closed polynomial form

$\kappa$5

with $\kappa$6 the Chebyshev polynomials of the second kind (Li et al., 26 May 2026).

This polynomial representation is especially useful because it gives explicit right and left eigenvectors away from EPs and an explicit Jordan construction at EPs. In the open chain, the eigenvectors split into two parity families with support on alternating sites, and the missing generalized eigenvectors at an EP are obtained by $\kappa$7-differentiation of the polynomial-branch eigenvector (Li et al., 26 May 2026). A plausible implication is that the non-Hermitian XY chain is one of the rare many-body settings where both ordinary diagonalization and defective-point Jordan structure remain analytically controlled.

3. Spectral reality, exceptional points, and non-Hermitian degeneracy

Exceptional points organize the spectral geometry of these chains. In the two-site alternating-field model,

$\kappa$8

with $\kappa$9 and $H_1$0. The EP occurs at $H_1$1, where $H_1$2 and $H_1$3 coalesce; for $H_1$4 they become purely imaginary. The $H_1$5-symmetric region is $H_1$6, while $H_1$7 is $H_1$8-broken (Li et al., 2022).

In the thermodynamic $H_1$9-symmetric chain, the unbroken–broken boundary is determined by

$h\to\lambda\in\mathbb C$0

which yields the analytic curve

$h\to\lambda\in\mathbb C$1

For finite $h\to\lambda\in\mathbb C$2, the phase boundary is staircase-like; in the thermodynamic limit it smooths into a hyperbola (Zhang et al., 2012). In the global complex-field model, by contrast, the condition $h\to\lambda\in\mathbb C$3 produces the critical ellipse

$h\to\lambda\in\mathbb C$4

which expands the Hermitian Ising critical points $h\to\lambda\in\mathbb C$5 into a critical transition zone (Liu et al., 2020).

Open-boundary, complex-anisotropy chains provide a more algebraic EP description. There, EPs are repeated roots of the same boundary polynomial,

$h\to\lambda\in\mathbb C$6

and the quasi-Hamiltonian becomes defective with Jordan blocks

$h\to\lambda\in\mathbb C$7

Near such points, the quasi-energies show square-root splitting,

$h\to\lambda\in\mathbb C$8

and a single loop around $h\to\lambda\in\mathbb C$9 permutes both eigenvalues and eigenstates (Li et al., 26 May 2026).

A distinct global picture emerges when the anisotropy parameter $H_2$0 itself is extended to complex values. For finite open chains, the EPs organize into two concentric rings in the complex $H_2$1-plane; for $H_2$2, $H_2$3, and in the $H_2$4 limit the rings collapse onto the unit circle $H_2$5. The same analysis identifies a broken $H_2$6-symmetric line along the pure imaginary $H_2$7-axis, with exactly four EPs on that line when $H_2$8 is a multiple of $H_2$9 (Henry et al., 6 Jul 2025).

4. Quantum phases and critical behavior

The phase structure of non-Hermitian XY chains depends strongly on how non-Hermiticity is introduced. In the complex global-field chain, three regions are distinguished by the non-Hermitian gap $\gamma\to i\kappa$0: a paramagnetic phase for $\gamma\to i\kappa$1, a ferromagnetic phase inside the critical ellipse, and a critical transition zone for $\gamma\to i\kappa$2 but outside the ellipse. The second derivatives of the ground-state energy density diverge on the ellipse, indicating a second-order quantum phase transition (Liu et al., 2020).

In the heralded non-Hermitian XY model with effective Hamiltonian

$\gamma\to i\kappa$3

the steady state is the right eigenstate with largest imaginary part, and the transition is controlled by the closing of

$\gamma\to i\kappa$4

The critical boundary is $\gamma\to i\kappa$5, separating short-range ordered and quasi-long-range ordered phases. The correlation-length exponent is $\gamma\to i\kappa$6 for $\gamma\to i\kappa$7 and $\gamma\to i\kappa$8 for $\gamma\to i\kappa$9, and the ordered phase exhibits frustrated spin patterns or a wavelength-4 spin-density wave depending on the couplings (Lee et al., 2014).

A major conceptual complication is that non-Hermitian quantum criticality is not uniquely defined by a single “ground state” or expectation-value prescription. For the two magnetic-field models analyzed in 2026, one may use the right eigenstate with minimal real-part energy and standard right-right expectation values,

$$H= -\frac{J}{2}\sum_{\ell=1}^{N}\bigl[(1+\gamma)\sigma_{\ell}^{x}\sigma_{\ell+1}^{x} +(1-\gamma)\sigma_{\ell}^{y}\sigma_{\ell+1}^{y}\bigr] -h\sum_{\ell}\sigma_{\ell}^{z} +i\eta\sum_{\ell}(-1)^{\ell}\sigma_{\ell}^{z},$$0

or the biorthogonal left-right prescription,

$$H= -\frac{J}{2}\sum_{\ell=1}^{N}\bigl[(1+\gamma)\sigma_{\ell}^{x}\sigma_{\ell+1}^{x} +(1-\gamma)\sigma_{\ell}^{y}\sigma_{\ell+1}^{y}\bigr] -h\sum_{\ell}\sigma_{\ell}^{z} +i\eta\sum_{\ell}(-1)^{\ell}\sigma_{\ell}^{z},$$1

The resulting phase diagram, magnetization, and long-distance correlations depend on both the formalism used and the state considered (Luo et al., 5 Jun 2026). In the complex-field model $$H= -\frac{J}{2}\sum_{\ell=1}^{N}\bigl[(1+\gamma)\sigma_{\ell}^{x}\sigma_{\ell+1}^{x} +(1-\gamma)\sigma_{\ell}^{y}\sigma_{\ell+1}^{y}\bigr] -h\sum_{\ell}\sigma_{\ell}^{z} +i\eta\sum_{\ell}(-1)^{\ell}\sigma_{\ell}^{z},$$2, the minimal-energy state yields ferromagnetic, Luttinger-liquid, and paramagnetic regions, while the steady state yields only Luttinger-liquid and paramagnetic behavior (Luo et al., 5 Jun 2026). This establishes that “the” phase diagram of a non-Hermitian XY chain is not unique without a preparation protocol.

For the alternating-field chain treated in two-spin cluster mean field, the many-body extension displays first-order magnetization jumps and concomitant jumps in concurrence at certain critical fields $$H= -\frac{J}{2}\sum_{\ell=1}^{N}\bigl[(1+\gamma)\sigma_{\ell}^{x}\sigma_{\ell+1}^{x} +(1-\gamma)\sigma_{\ell}^{y}\sigma_{\ell+1}^{y}\bigr] -h\sum_{\ell}\sigma_{\ell}^{z} +i\eta\sum_{\ell}(-1)^{\ell}\sigma_{\ell}^{z},$$3 whenever $$H= -\frac{J}{2}\sum_{\ell=1}^{N}\bigl[(1+\gamma)\sigma_{\ell}^{x}\sigma_{\ell+1}^{x} +(1-\gamma)\sigma_{\ell}^{y}\sigma_{\ell+1}^{y}\bigr] -h\sum_{\ell}\sigma_{\ell}^{z} +i\eta\sum_{\ell}(-1)^{\ell}\sigma_{\ell}^{z},$$4, with the discontinuities rounded at finite temperature (Li et al., 2022).

5. Correlations, concurrence, and the choice of observables

Entanglement is one of the most explicit diagnostics of non-Hermitian criticality in these chains. In the two-site alternating-field model, the Wootters concurrence of the nondegenerate ground state is

$$H= -\frac{J}{2}\sum_{\ell=1}^{N}\bigl[(1+\gamma)\sigma_{\ell}^{x}\sigma_{\ell+1}^{x} +(1-\gamma)\sigma_{\ell}^{y}\sigma_{\ell+1}^{y}\bigr] -h\sum_{\ell}\sigma_{\ell}^{z} +i\eta\sum_{\ell}(-1)^{\ell}\sigma_{\ell}^{z},$$5

when $$H= -\frac{J}{2}\sum_{\ell=1}^{N}\bigl[(1+\gamma)\sigma_{\ell}^{x}\sigma_{\ell+1}^{x} +(1-\gamma)\sigma_{\ell}^{y}\sigma_{\ell+1}^{y}\bigr] -h\sum_{\ell}\sigma_{\ell}^{z} +i\eta\sum_{\ell}(-1)^{\ell}\sigma_{\ell}^{z},$$6, while

$$H= -\frac{J}{2}\sum_{\ell=1}^{N}\bigl[(1+\gamma)\sigma_{\ell}^{x}\sigma_{\ell+1}^{x} +(1-\gamma)\sigma_{\ell}^{y}\sigma_{\ell+1}^{y}\bigr] -h\sum_{\ell}\sigma_{\ell}^{z} +i\eta\sum_{\ell}(-1)^{\ell}\sigma_{\ell}^{z},$$7

when $$H= -\frac{J}{2}\sum_{\ell=1}^{N}\bigl[(1+\gamma)\sigma_{\ell}^{x}\sigma_{\ell+1}^{x} +(1-\gamma)\sigma_{\ell}^{y}\sigma_{\ell+1}^{y}\bigr] -h\sum_{\ell}\sigma_{\ell}^{z} +i\eta\sum_{\ell}(-1)^{\ell}\sigma_{\ell}^{z},$$8. In the $$H= -\frac{J}{2}\sum_{\ell=1}^{N}\bigl[(1+\gamma)\sigma_{\ell}^{x}\sigma_{\ell+1}^{x} +(1-\gamma)\sigma_{\ell}^{y}\sigma_{\ell+1}^{y}\bigr] -h\sum_{\ell}\sigma_{\ell}^{z} +i\eta\sum_{\ell}(-1)^{\ell}\sigma_{\ell}^{z},$$9-symmetric region, the concurrence of the $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$00 ground state is maximally entangled and independent of $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$01 and $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$02; in the $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$03-broken region, it decays for $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$04 and has a cusp at the EP $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$05. The same non-analyticity persists in the biorthogonal basis, where the concurrence drops at the EP (Li et al., 2022).

Thermal entanglement is similarly nontrivial. In the isotropic limit $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$06, the imaginary field $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$07 weakens thermal entanglement and lowers the sudden-death temperature. In the Ising limit $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$08, a finite $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$09 enhances thermal entanglement and raises the critical temperature below which $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$10. For intermediate $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$11, the dependence of $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$12 on $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$13 is non-monotonic, with a minimum at the degeneracy line $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$14 (Li et al., 2022).

For many-body chains, correlation functions rather than two-qubit entanglement become central. In the global complex-field model, the long-distance behavior of

$$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$15

distinguishes the phases: exponential decay in the paramagnet, constant asymptote in the ferromagnet, and power-law decay in the critical transition zone (Liu et al., 2020). In the 2026 comparison of formalisms, the long-distance asymptotics of $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$16 and $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$17 differ qualitatively between RR and BO/LR prescriptions; for example, in the imaginary-anisotropy model $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$18, BO correlators become singular or ill-defined at $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$19, whereas RR observables remain real and physically sensible (Luo et al., 5 Jun 2026).

Biorthogonal fidelity and entanglement scaling provide an additional layer of characterization. For non-Hermitian XY extensions in a magnetic field, the peak of the biorthogonal fidelity susceptibility obeys

$$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$20

recovering $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$21 for Ising-type transitions and $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$22 for the anisotropy transition in the complex-anisotropy model. The biorthogonal entanglement entropy satisfies

$$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$23

with $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$24 numerically on the anisotropy line of $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$25 (Liu et al., 23 Jan 2025). The same work finds that the entanglement transition goes hand in hand with the non-Hermitian topological phase transition (Liu et al., 23 Jan 2025).

6. Topology, frustration, and boundary effects

Boundary conditions are not a technical detail in non-Hermitian XY chains; they can change the ground-state sector and the topological interpretation. In the odd-length ring-frustrated chain, Jordan–Wigner transformation produces parity-dependent boundary conditions because the boundary term carries a sign $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$26. In the real-spectrum kink phase,

$$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$27

the true spin-chain ground state is not the Bogoliubov vacuum but the one-mode-occupied state

$$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$28

Its low-energy excitations are kink–antikink modes, and the phase is gapless with topological invariant $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$29; by contrast, the real-spectrum paramagnetic phase at $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$30 has $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$31 (Bi et al., 2020).

Topological diagnostics also arise directly from the BdG structure. For the complex-field and complex-anisotropy models, the non-Hermitian Bloch Hamiltonian anticommutes with $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$32, allowing the definition of a complex angle

$$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$33

The winding number remains integer-valued and distinguishes trivial $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$34 from non-trivial $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$35 phases (Liu et al., 23 Jan 2025). In the open-chain $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$36-model, the thermodynamic EP rings collapse onto $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$37, exactly the boundary between the $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$38 and $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$39 topological sectors (Henry et al., 6 Jul 2025).

The open-boundary exact solution with complex anisotropy sharpens the topological interpretation of EPs. Because the eigenstates are algebraic functions of $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$40, the branch-cut structure of the biorthogonal eigenstates directly shows the exchange of eigenstates when an EP is encircled, and the permutation of single-particle quasi-energies lifts to the permutation of many-body levels (Li et al., 26 May 2026). This suggests that the non-Hermitian XY chain is a particularly transparent real-space platform for many-body EP topology beyond momentum-space descriptions.

7. Hermitian counterparts, disorder, and physical realization

Several works ask to what extent a non-Hermitian XY chain can be related to a Hermitian one. In the dimerized chain with alternating imaginary field, a full real spectrum can appear only in presence of dimerization. For $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$41, one can construct an equivalent Hermitian XY chain by renormalizing the couplings: $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$42 and, in the anisotropic case,

$$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$43

with an adjusted field $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$44 so that the Hermitian chain reproduces the non-Hermitian spectrum and phase boundaries (Giorgi, 2010). In the $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$45-symmetric chain, a Hermitian counterpart with identical real spectrum can also be constructed in the biorthogonal basis, though it is generally non-local; far from the exceptional curve it reduces approximately to an ordinary isotropic XY chain in a transverse field (Zhang et al., 2012).

The physical origin of non-Hermiticity varies. The complex-field model $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$46 can be derived as the no-jump effective Hamiltonian of a Lindblad master equation with loss operators $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$47 and rate $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$48 (Luo et al., 5 Jun 2026). The heralded magnetism proposal realizes the effective Hamiltonian through three-level atoms with spontaneous decay and post-selection on null photon records, with suggested platforms including trapped ions, cavity QED, and atoms in optical lattices (Lee et al., 2014). The $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$49-symmetric model has been connected to arrays of coupled optical waveguides with balanced gain and loss, superconducting-qubit chains with engineered dissipation, and ultracold atoms with controlled complex tunneling amplitudes (Zhang et al., 2012).

Exact solvability is fragile under generic perturbations. When a random field $$H \;=\;\sum_{j=1}^N\Bigl[ \tfrac{1+\gamma}{2}\,\sigma_j^x\sigma_{j+1}^x +\tfrac{1-\gamma}{2}\,\sigma_j^y\sigma_{j+1}^y -h\,\sigma_j^z \;+\;i\,\kappa\,\sigma_j^x\sigma_{j+1}^y \Bigr],$$50 is added to the non-Hermitian anisotropic XY chain, translational invariance is broken and the free-fermion solvability is spoiled, producing a non-integrable model. Complex spacing ratios then show a crossover from Poisson-like to Ginibre-unitary behavior, and the phenomenology is compared directly with one-parameter random-matrix interpolations between 1D-Poisson or 2D-Poisson statistics and GinUE (Sarkar et al., 2023). This places the non-Hermitian XY chain at an intersection of integrable many-body theory, non-Hermitian symmetry breaking, and spectral quantum chaos.

Overall, the non-Hermitian XY spin chain is not a single model but a mathematically linked family. Across staggered imaginary fields, global complex transverse fields, imaginary anisotropies, odd-ring frustration, and open-boundary complex anisotropy, the common structure is a free-fermion backbone enriched by anti-linear symmetry, EP singularity, competing prescriptions for observables, and topological reorganization of the spectrum (Li et al., 2022, Luo et al., 5 Jun 2026).