Higher-Order Kuramoto Model Dynamics

• Higher-Order Kuramoto Model is a generalization of the classic Kuramoto framework, incorporating higher Fourier harmonics and many-body interactions.
• It employs formulations on hypergraphs and simplicial complexes to reveal novel collective states such as clustering, multistability, and twisted synchronization.
• Analytical reductions like Ott–Antonsen and Watanabe–Strogatz provide insights into mean-field dynamics and control strategies for managing complex synchronization.

The higher-order Kuramoto model is a family of phase-oscillator systems that extends the classical Kuramoto coupling law beyond the pairwise first harmonic $\sin(\theta_j-\theta_i)$. In this family, “higher-order” can denote a single higher Fourier harmonic, such as $\sin[q(\theta_i-\theta_j)]$, genuine nonpairwise many-body terms such as $\sin(2\theta_j-\theta_k-\theta_i)$ or $\sin(\theta_{j_1}+\cdots+\theta_{j_n}-n\theta_i)$, or dynamics defined on links, triangles, and higher-dimensional simplices rather than only on nodes (Delabays, 2019, Suman et al., 2024, León et al., 27 Mar 2025, Millán et al., 2019). A central distinction is between the simple $q$th-order harmonic model, which is dynamically equivalent to the standard Kuramoto model under a linear covering transformation, and models with multiple harmonics or genuine many-body couplings, where clustering, multistability, explosive synchronization, twisted states, and richer bifurcation scenarios arise (Delabays, 2019, Bick et al., 2022, Costa et al., 2024).

1. Canonical formulations

There is no single canonical higher-order Kuramoto equation. The literature contains several structurally distinct generalizations, each emphasizing a different mechanism by which nonclassical collective behavior enters.

Formulation	Representative dynamics	Interpretation
Simple $q$th-order harmonic	$$\dot\theta_i=\omega_i-\frac{K}{N}\sum_j \sin[q(\theta_i-\theta_j)]$$	Single Fourier harmonic; $q$-branch clustering
Pairwise plus triadic	$$\dot\theta_i=\omega_i+\frac{K_1}{N}\sum_j\sin(\theta_j-\theta_i)+\frac{K_2}{N^2}\sum_{j,k}\sin(2\theta_j-\theta_k-\theta_i)$$	First nontrivial multi-body term
Hypergraph $p$-body model	$\sin[q(\theta_i-\theta_j)]$0	Interactions on hyperedges of size $\sin[q(\theta_i-\theta_j)]$1

These representative forms appear, respectively, in the single-harmonic model, the globally coupled triadic extension, and the hypergraph formulation with interactions of all orders (Delabays, 2019, Suman et al., 2024, Moriamé et al., 2024).

A distinct simplicial-complex formulation places oscillators on $\sin[q(\theta_i-\theta_j)]$2-simplices and couples them through incidence matrices $\sin[q(\theta_i-\theta_j)]$3 and $\sin[q(\theta_i-\theta_j)]$4: $\sin[q(\theta_i-\theta_j)]$5 In that setting, the relevant geometric objects are the higher-order Laplacians

$\sin[q(\theta_i-\theta_j)]$6

together with the Hodge decomposition into harmonic, irrotational, and solenoidal components (Millán et al., 2019).

A more general route starts from weakly coupled limit-cycle oscillators on hypergraphs or simplicial complexes,

$\sin[q(\theta_i-\theta_j)]$7

and uses phase reduction to obtain a skeleton phase model

$\sin[q(\theta_i-\theta_j)]$8

In this formulation the higher-order topology is preserved at first order, and under odd symmetry $\sin[q(\theta_i-\theta_j)]$9, only odd couplings survive at $\sin(2\theta_j-\theta_k-\theta_i)$0 (León et al., 27 Mar 2025).

2. Exact equivalence and low-dimensional reductions

For the simple $\sin(2\theta_j-\theta_k-\theta_i)$1th-order harmonic model,

$\sin(2\theta_j-\theta_k-\theta_i)$2

the change of variables $\sin(2\theta_j-\theta_k-\theta_i)$3 gives

$\sin(2\theta_j-\theta_k-\theta_i)$4

which is exactly the standard Kuramoto equation in the $\sin(2\theta_j-\theta_k-\theta_i)$5-variables. The corresponding complex order parameter

$\sin(2\theta_j-\theta_k-\theta_i)$6

becomes the usual first-order parameter in the lifted coordinates, so $\sin(2\theta_j-\theta_k-\theta_i)$7. In this sense, all synchronization thresholds, stability properties, basin-size estimates and noise-driven escape rates are carried over by the rescalings $\sin(2\theta_j-\theta_k-\theta_i)$8, $\sin(2\theta_j-\theta_k-\theta_i)$9, and $\sin(\theta_{j_1}+\cdots+\theta_{j_n}-n\theta_i)$0, while the $\sin(\theta_{j_1}+\cdots+\theta_{j_n}-n\theta_i)$1 different lifts of a synchronous solution account for clustering into $\sin(\theta_{j_1}+\cdots+\theta_{j_n}-n\theta_i)$2 branches. The equivalence extends trivially to weighted graph coupling, but it fails when more than one Fourier mode is present (Delabays, 2019).

For globally coupled models with pairwise and triadic interactions, Ott–Antonsen reduction yields a one-dimensional amplitude flow in the thermodynamic limit. With Lorentzian frequency width $\sin(\theta_{j_1}+\cdots+\theta_{j_n}-n\theta_i)$3,

$\sin(\theta_{j_1}+\cdots+\theta_{j_n}-n\theta_i)$4

which already displays the coexistence of incoherent, unstable, and synchronized fixed points in appropriate parameter ranges (Suman et al., 2024). A broader asymmetric arbitrary-order class admits an exact Ott–Antonsen reduction in which the macroscopic driving collapses to

$\sin(\theta_{j_1}+\cdots+\theta_{j_n}-n\theta_i)$5

and the order-parameter modulus obeys

$\sin(\theta_{j_1}+\cdots+\theta_{j_n}-n\theta_i)$6

In that framework, the contribution of a term depends not on the actual simplex size, but on the “effective order”

$\sin(\theta_{j_1}+\cdots+\theta_{j_n}-n\theta_i)$7

which organizes the macroscopic dynamics (Costa et al., 10 Jan 2025).

For identical oscillators, Watanabe–Strogatz theory supplies an exact finite-$\sin(\theta_{j_1}+\cdots+\theta_{j_n}-n\theta_i)$8 reduction. For pure $\sin(\theta_{j_1}+\cdots+\theta_{j_n}-n\theta_i)$9th-harmonic coupling, the system has three effective collective degrees of freedom $q$0 plus $q$1 constants of motion (Gong et al., 2019). A later WS formulation covers a broad class of pairwise and higher-order globally coupled models and shows that the dynamics of the WS parameters coincide with those of the mean-field parameters; the poles of the associated Möbius transformation act as basin boundaries for both global and cluster synchronization (Jain et al., 19 Aug 2025).

3. Collective states, synchronization transitions, and spectral structure

Higher-order coupling changes the set of admissible collective states. In the simple $q$2th-order harmonic model, large $q$3 signifies that the phases lie in at most $q$4 tight clusters separated by approximately $q$5 (Delabays, 2019). In the pairwise-plus-symmetric-triadic model,

$q$6

the higher-order term favors two-cluster states. A self-consistent mean-field treatment introduces a clustering order parameter $q$7, distinguishes fully synchronized, uniform-cluster, bimodal-cluster, standing-wave, and incoherent regimes, and identifies a $q$8-transition: when the cluster separation reaches $q$9, the locked cluster state destabilizes abruptly (Carballosa et al., 2023).

Bistability and abrupt synchronization are recurrent motifs. In the triadic globally coupled model reduced by Ott–Antonsen, the interval $q$0 contains three positive roots of the amplitude equation: a stable incoherent fixed point $q$1, an unstable saddle $q$2, and a stable synchronized state $q$3 (Suman et al., 2024). On simplicial complexes, an adaptive coupling in which the global curl and divergence order parameters feed into one another produces explosive synchronization with hysteresis, whereas the corresponding uncoupled projected systems show continuous onset and no hysteresis (Millán et al., 2019).

The spectral theory of higher-order models is correspondingly richer. In the pure three-body globally coupled model,

$q$4

the interaction can be rewritten as $q$5, and the continuum linearization takes the form $q$6, where $q$7 is rank two. If the natural-frequency distribution has infinite support, drifting oscillators generate a continuous spectrum on the imaginary axis, so stationary multicluster states are at best neutrally stable. If the distribution has finite support and all oscillators are phase-locked, the spectrum becomes real and fully locked branches can be linearly stable or unstable, with abrupt desynchronization at a saddle-node point $q$8 (Xu et al., 2020).

In nonlocal ring geometries, higher-order interactions stabilize patterned states unavailable to the classical model. For identical oscillators on an infinite ring with pairwise, triplet, and quadruplet interactions, uniformly $q$9-twisted states

$$\dot\theta_i=\omega_i-\frac{K}{N}\sum_j \sin[q(\theta_i-\theta_j)]$$0

have Fourier-diagonal linearization with eigenvalues

$$\dot\theta_i=\omega_i-\frac{K}{N}\sum_j \sin[q(\theta_i-\theta_j)]$$1

Triplet and quadruplet couplings shift the classical stability windows and can stabilize twisted states that are unstable under pairwise coupling alone; the resulting pitchfork bifurcations may be supercritical or subcritical depending on the higher-order weights (Bick et al., 2022).

4. Topology, heterogeneity, and mean-field theory

A major theme in recent work is that higher-order topology survives phase reduction. In the general hypergraph-to-simplicial-complex reduction, adjacency tensors and boundary operators enter the reduced phase model explicitly, so the phase equations inherit the original higher-order topology rather than replacing it by a pairwise approximation (León et al., 27 Mar 2025). For odd-symmetric oscillators, even-$$\dot\theta_i=\omega_i-\frac{K}{N}\sum_j \sin[q(\theta_i-\theta_j)]$$2 interactions vanish at $$\dot\theta_i=\omega_i-\frac{K}{N}\sum_j \sin[q(\theta_i-\theta_j)]$$3, and the surviving terms can be written in Kuramoto-like sine form

$$\dot\theta_i=\omega_i-\frac{K}{N}\sum_j \sin[q(\theta_i-\theta_j)]$$4

providing a systematic derivation of higher-order Kuramoto families from microscopic oscillator dynamics (León et al., 27 Mar 2025).

A rigorous multi-population framework is available at the mean-field level. For $$\dot\theta_i=\omega_i-\frac{K}{N}\sum_j \sin[q(\theta_i-\theta_j)]$$5 interacting populations with general $$\dot\theta_i=\omega_i-\frac{K}{N}\sum_j \sin[q(\theta_i-\theta_j)]$$6-body couplings, the empirical measures $$\dot\theta_i=\omega_i-\frac{K}{N}\sum_j \sin[q(\theta_i-\theta_j)]$$7 satisfy transport equations

$$\dot\theta_i=\omega_i-\frac{K}{N}\sum_j \sin[q(\theta_i-\theta_j)]$$8

with existence and uniqueness proved in the bounded-Lipschitz/Wasserstein setting. The all-synchronized manifold $$\dot\theta_i=\omega_i-\frac{K}{N}\sum_j \sin[q(\theta_i-\theta_j)]$$9 and the all-splay manifold $q$0 are invariant; the all-synchronized state is never asymptotically stable in the full generality considered, but under $q$1 assumptions and positivity of the local phase-derivative coefficients $q$2, it is Lyapunov stable. In a sinusoidal higher-order example of Skardal–Arenas type, the stability threshold becomes $q$3, so nonpairwise coupling shifts the classical criterion (Bick et al., 2020).

For finite hypergraphs with dyadic and triadic couplings, a self-consistent analytical hierarchy has recently been developed using generalized local order parameters

$q$4

This yields Time-Averaged Theory, a Frequency-Distribution Approximation, and a Heterogeneous Mean-Field closure. The continuous synchronization threshold is determined by the Perron–Frobenius eigenvalue $q$5 of the dyadic adjacency matrix,

$q$6

while the onset of bistability occurs at

$q$7

showing that the critical triadic coupling depends on correlations between the leading dyadic eigenvectors and the triadic interaction structure (Kumpeerakij et al., 23 May 2026).

5. Finite-size, forcing, delay, and stochastic effects

Finite-size effects are not perturbative decorations of the thermodynamic picture; in some regimes they qualitatively alter the macroscopic dynamics. In the globally coupled model with triadic interactions, finite-$q$8 fluctuations satisfy $q$9. Within the bistable window, these fluctuations induce rare escapes across the unstable saddle, producing a first-exit-time distribution

$$\dot\theta_i=\omega_i+\frac{K_1}{N}\sum_j\sin(\theta_j-\theta_i)+\frac{K_2}{N^2}\sum_{j,k}\sin(2\theta_j-\theta_k-\theta_i)$$0

and a synchronization probability that shifts to smaller $$\dot\theta_i=\omega_i+\frac{K_1}{N}\sum_j\sin(\theta_j-\theta_i)+\frac{K_2}{N^2}\sum_{j,k}\sin(2\theta_j-\theta_k-\theta_i)$$1 as $$\dot\theta_i=\omega_i+\frac{K_1}{N}\sum_j\sin(\theta_j-\theta_i)+\frac{K_2}{N^2}\sum_{j,k}\sin(2\theta_j-\theta_k-\theta_i)$$2 decreases. In a two-population adaptive extension, finite size creates an additional partially synchronized attractor in the $$\dot\theta_i=\omega_i+\frac{K_1}{N}\sum_j\sin(\theta_j-\theta_i)+\frac{K_2}{N^2}\sum_{j,k}\sin(2\theta_j-\theta_k-\theta_i)$$3-plane that disappears as $$\dot\theta_i=\omega_i+\frac{K_1}{N}\sum_j\sin(\theta_j-\theta_i)+\frac{K_2}{N^2}\sum_{j,k}\sin(2\theta_j-\theta_k-\theta_i)$$4 (Suman et al., 2024).

External forcing combined with higher-order interactions yields a notably large bifurcation atlas. For a forced model with pairwise, three-body, and four-body terms, Ott–Antonsen reduction closes the mean field to

$$\dot\theta_i=\omega_i+\frac{K_1}{N}\sum_j\sin(\theta_j-\theta_i)+\frac{K_2}{N^2}\sum_{j,k}\sin(2\theta_j-\theta_k-\theta_i)$$5

$$\dot\theta_i=\omega_i+\frac{K_1}{N}\sum_j\sin(\theta_j-\theta_i)+\frac{K_2}{N^2}\sum_{j,k}\sin(2\theta_j-\theta_k-\theta_i)$$6

with $$\dot\theta_i=\omega_i+\frac{K_1}{N}\sum_j\sin(\theta_j-\theta_i)+\frac{K_2}{N^2}\sum_{j,k}\sin(2\theta_j-\theta_k-\theta_i)$$7. When the unforced system is bistable, saddle-node, Hopf, SNIPER, and homoclinic manifolds are duplicated, partitioning the $$\dot\theta_i=\omega_i+\frac{K_1}{N}\sum_j\sin(\theta_j-\theta_i)+\frac{K_2}{N^2}\sum_{j,k}\sin(2\theta_j-\theta_k-\theta_i)$$8-plane into 11 asymptotic regions that include competing forced and spontaneous synchronization (Costa et al., 2024).

Time delay can itself generate effective higher-order interactions. Starting from the delayed Kuramoto model

$$\dot\theta_i=\omega_i+\frac{K_1}{N}\sum_j\sin(\theta_j-\theta_i)+\frac{K_2}{N^2}\sum_{j,k}\sin(2\theta_j-\theta_k-\theta_i)$$9

a weak-coupling, small-delay expansion to $p$0 produces an effective delay-free model with a Kuramoto–Sakaguchi pairwise term and a genuine three-body interaction. In the identical-oscillator limit,

$p$1

Ott–Antonsen reduction then yields an amplitude equation $p$2, and the resulting stability diagram for incoherent and synchronized states closely matches the original delayed system for $p$3 and $p$4 (Fujii et al., 18 Dec 2025).

Under non-Gaussian Lévy noise, synchronization boundaries shift substantially. In the stochastic higher-order Kuramoto model with pairwise and triadic interactions, lower stability index $p$5 weakens synchronization, stronger coupling is required than under Gaussian white noise, sufficiently large noise can remove bistability, and spike statistics display power-law tails in large inter-window intervals together with long-memory signatures in generalized spectral analysis (Zhao et al., 29 Sep 2025).

6. Synthesis from microscopic oscillators, control, and functional uses

Higher-order Kuramoto equations are not restricted to phenomenological phase models. For arbitrary smooth limit-cycle oscillators, specially designed pairwise and three-body interaction functions built from the asymptotic phase $p$6 and phase-sensitivity function $p$7 can be chosen so that phase reduction yields exactly the desired higher-order Kuramoto model at $p$8. In particular, interaction terms of the form

$p$9

and

$\sin[q(\theta_i-\theta_j)]$00

produce purely sinusoidal first- and higher-order phase-coupling functions. Numerical validation with FitzHugh–Nagumo oscillators shows close agreement between the full oscillator dynamics and the reduced higher-order Kuramoto description, and an OA-based control $\sin[q(\theta_i-\theta_j)]$01 stabilizes the collective relative phase $\sin[q(\theta_i-\theta_j)]$02 at $\sin[q(\theta_i-\theta_j)]$03 (Namura et al., 16 Oct 2025).

Higher-order synchronization can also be actively suppressed. A Hamiltonian embedding of the higher-order Kuramoto model introduces conjugate action-angle variables $\sin[q(\theta_i-\theta_j)]$04 and an invariant torus $\sin[q(\theta_i-\theta_j)]$05 on which the phase dynamics reproduce the original higher-order system. Hamiltonian control theory then constructs a feedback perturbation

$\sin[q(\theta_i-\theta_j)]$06

whose restriction to the invariant torus yields desynchronizing feedback terms $\sin[q(\theta_i-\theta_j)]$07. Numerical experiments on all-to-all hypergraphs, random simplicial complexes, and an empirical cat-cortex hypergraph show that the full higher-order control suppresses synchronization across wide $\sin[q(\theta_i-\theta_j)]$08 domains; pairwise-only control is often sufficient when $\sin[q(\theta_i-\theta_j)]$09 is not too small relative to $\sin[q(\theta_i-\theta_j)]$10, and dense-hypergraph pinning requires a controlled fraction near $\sin[q(\theta_i-\theta_j)]$11 for full desynchronization (Moriamé et al., 2024).

A different functional direction uses higher-order couplings to build dense associative memory. In a generalized Kuramoto network with pairwise second-harmonic and quartic fourth-harmonic Hebbian couplings,

$\sin[q(\theta_i-\theta_j)]$12

mean-field theory yields a tricritical point

$\sin[q(\theta_i-\theta_j)]$13

separating continuous from discontinuous retrieval. In the quartic-dominated regime, memory and incoherent states coexist across a bistable window, and Kramers escape from a memory state obeys $\sin[q(\theta_i-\theta_j)]$14 (Nagerl et al., 29 Jul 2025).

Higher-order interactions are not uniformly synchrony-impairing. Numerical analysis on random hypergraphs shows that strong higher-order coupling generally works against synchronization from incoherent initial data, but weak higher-order coupling can enhance synchronization, and under a fixed budget of pairwise and higher-order interactions, mixed allocations consistently outperform purely pairwise or purely higher-order architectures (Muolo et al., 14 Aug 2025). Taken together with the exact equivalence results for simple harmonic lifting, the phase-reduction constructions, and the control frameworks, this establishes the higher-order Kuramoto model as a broad theoretical class rather than a single equation: some members are reducible to the classical Kuramoto model, while others support genuinely new topology-dependent and many-body collective dynamics.