Colossal topological Hall effect at the transition between isolated and lattice-phase interfacial skyrmions

Abstract

The topological Hall effect is used extensively to study chiral spin textures in various materials. However, the factors controlling its magnitude in technologically-relevant thin films remain uncertain. Using variable-temperature magnetotransport and real-space magnetic imaging in a series of Ir/Fe/Co/Pt heterostructures, here we report that the chiral spin fluctuations at the phase boundary between isolated skyrmions and a disordered skyrmion lattice result in a power-law enhancement of the topological Hall resistivity by up to three orders of magnitude. Our work reveals the dominant role of skyrmion stability and configuration in determining the magnitude of the topological Hall effect.

Introduction

Magnetic skyrmions are topologically charged nanoscale spin textures that form due to competition between spin-rotating and spin-aligning magnetic interactions. In thin-film heterostructures, these magnetic interactions can be finely tuned via the multilayer geometry and composition, rendering skyrmion-hosting films ideally suited for spintronic applications^1,2,3,4,5. One promising route towards functionalizing skyrmions in metallic systems is to utilize their intrinsic magnetoelectric coupling, which is manifested by a topological Hall effect (THE)^6,7,8. Charge carriers moving through a skyrmion spin texture experience an emergent magnetic field (B_eff) associated with the spin winding of a skyrmion. The transverse deflection of charge carriers interacting with B_eff results in the THE⁷. Provided that each skyrmion remains stationary with respect to incident charge carriers, an array of skyrmions will exhibit a topological Hall resistivity ρ^THE, given by

$${\rho }^{\mathrm{THE}}=P\cdot {R}_{0}^{\prime}\cdot ({n}_{\mathrm{sk}}\cdot {{{\Phi }}}_{0}),$$

(1)

where P is the spin polarization of the charge carriers, ${R}_{0}^{\prime}$ is the Hall coefficient representing the effective charge density contributing to the THE (usually taken as the classical Hall coefficient R₀), B_eff ≡ n_sk ⋅ Φ₀ is the emergent field associated with a given skyrmion density n_sk, and Φ₀ = h/e is the magnetic flux quantum, with h Planck’s constant and −e the electron charge. This phenomenon is distinct from the classical and anomalous Hall effects, which are proportional to the applied magnetic field H and magnetization M(H), respectively⁷. Experimentally, ρ^THE(H) can be identified as the residual between the total measured Hall resistivity ρ_yx(H) and a fit to the classical R₀(H) and anomalous R_SM(H) Hall resistivities, namely ${\rho }_{yx}^{\mathrm{fit}}(H)={R}_{0}H+{R}_{{\rm{S}}}M(H)$⁹ (where R₀ and R_S are the classical and anomalous Hall coefficients, respectively). The anomalous Hall resistivity ρ^AHE ≡ R_SM(H) can be described by a superposition of terms with linear and quadratic dependences on the longitudinal resistivity ρ_xx, corresponding to the skew scattering and side jump terms, respectively⁸. However, the variation of ρ_xx(H) in these multilayers is extremely small (<0.01% for fields up to magnetic saturation at H_S and <0.17% for fields up to ±5 T^5,9). Consequently, the treatment of ρ^AHE as $[a\cdot {\rho }_{xx}(H)+b\cdot {\rho }_{xx}^{2}(H)]M(H)$ or simply R_SM(H) has negligible influence on the estimated ρ^THE9. A detailed scaling analysis of ρ^AHE with ρ_xx as a function of temperature T (5–300 K) and H for these multilayers can be found in refs. ^5,9, together with a discussion of the validity and reproducibility of the estimated ρ^THE.

Using Eq. (1) and our experimentally-determined ρ^THE, one can hence estimate n_sk from an electrical transport measurement as:

$${n}_{\mathrm{sk}}({\mathrm{THE}})=| {\rho }^{\mathrm{THE}}|\,\div\,|(P\cdot {R}_{0}\cdot {{{\Phi }}}_{0})| .$$

(2)

n_sk may also be measured directly using real-space imaging techniques such as magnetic force microscopy (MFM)^9,10,11, magnetic transmission X-ray microscopy^4,5, or Lorentz transmission electron microscopy¹². Comparing these transport and imaging approaches can yield evidence for adiabatic transport if n_sk(THE) ≈ n_sk(MFM)^6,7,8, non-adiabaticity if n_sk(THE) < n_sk(MFM)^13,14, or alternatively reveal enhanced transverse scattering mechanisms if n_sk(THE) > n_sk(MFM)^13,15,16,17.

In bulk non-centrosymmetric materials which exhibit stable skyrmion lattices, the values of n_sk estimated from transport and imaging are in good agreement^6,7,8. However, this is not the case for thin-film multilayers with an interfacial Dzyaloshinskii–Moriya interaction (DMI), which can be tuned to exhibit isolated or dense skyrmion configurations. Large, conflicting values for ρ^THE have been reported in magnetic thin films, corresponding to n_sk(THE) orders of magnitude larger than n_sk(MFM)^{5,10,11,12,18,19}.

Determining the mechanism leading to this extraordinary disagreement is crucial for understanding the electrical response of chiral spin textures and their detection in devices. Recently, a valuable clue has been provided by predictions^15,16,17 and observations^20,21 of spin fluctuation-induced effects on charge transport in non-coplanar magnets. It is hence plausible that quantum or thermal fluctuations may influence the Hall response of materials with a chiral instability.

In this work, using temperature-dependent Hall transport and MFM, we track the evolution of the THE across the transition between isolated skyrmions and a disordered skyrmion lattice. We find that isolated skyrmion configurations produce a larger THE than dense arrays of skyrmions, with an enhancement of up to three orders of magnitude at the transition. Our data reveal a power-law behavior in n_sk(THE)/n_sk(MFM), which we interpret in terms of chiral spin fluctuations. Following universal scaling laws, we extract the critical exponents governing this phase transition^22,23,24.

Results

Figure 1 shows measurements of ρ^THE(H) at 300 K for a set of Ir/Fe(x)/Co(y)/Pt multilayers. MFM images acquired at H maximizing ρ^THE(H) display spin configurations ranging from isolated skyrmions (Fig. 1a–c) to dense, disordered skyrmion lattices (Fig. 1d–g) (see Section 1 of the Supplementary Information for the evolution of magnetic textures). This evolution in skyrmion configuration is driven by the T-dependent stability parameter $\kappa \equiv \,\pi D/4\sqrt{A{K}_{\mathrm{eff}}}$, which describes the competition between the three key magnetic interactions: exchange coupling (A), magnetic anisotropy (K_eff), and DMI (D). Our multilayer Ir/Fe(x)/Co(y)/Pt stacks allow us to systematically tune κ (and hence n_sk) via the Fe/Co layer thickness ratio: x/y < 0.5 yields κ < 1, whereas κ ≥ 1 for x/y ≥ 0.5⁵. Varying T provides an additional handle to tune κ: for a given Fe/Co ratio, κ increases with T due to T-dependent magnetic interactions (see Section 2 of the Supplementary Information). At H = 0, these chiral magnetic films exhibit a labyrinthine stripe domain phase. Under finite H, this transforms into a metastable skyrmion phase if κ < 1 (Fig. 1a–c), or a disordered skyrmion lattice phase if κ ≥ 1 (Fig. 1d–f). For κ ≥ 1, the skyrmion lattice dissolves into isolated skyrmions before a uniformly polarized ferromagnetic (FM) phase develops for H > H_S. The isolated skyrmion phase emerges between a lattice and a FM phase; its appearance can be regulated by κ, T, H, or a combination of these parameters (see Section 1 of the Supplementary Information). Indeed, in chiral magnetic films, the transformation of a polarized FM phase into an array of isolated skyrmions and ultimately a skyrmion lattice occurs via a nucleation-type second-order phase transition^22,23,24.

Fig. 1: Enhanced topological Hall resistivity (ρ^THE) for isolated skyrmions compared with dense skyrmion arrays.

a–c ρ^THE(H) of isolated skyrmion configurations observed in [Fe2/Co8]₂₀, [Fe2/Co6]₂₀, and [Fe2/Co5]₂₀. d–f ρ^THE(H) of dense skyrmion arrays observed in [Fe5/Co5]₂₀, [Fe4/Co6]₂₀, and [Fe4/Co4]₂₀. g Δ_ρ/M ≡ n_sk(THE)/n_sk(MFM) estimated from transport and imaging experiments shown in (a–f) and in Section 1 of the Supplementary Information. n_sk(THE) is given by ρ^THE/(P ⋅ R₀ ⋅ Φ₀) with P = 0.56 and n_sk(MFM) is estimated from MFM images; details of our image analysis methods can be found in our earlier reports^{5, 9, 30}. The detailed evolution of magnetic textures with H can be found in Section 1 of the Supplementary Information. Scale bars in white (MFM images corresponding to (a–e)) are 1 μm and in black (MFM images corresponding to (f)) 0.5 μm. All the images are acquired at a scan height of 20 nm. The color scale represents the phase shift in the MFM signal due to the magnetic force acting on the tip. The ρ^THE(H) profiles show a small non-zero offset ≤2 nΩ cm above the saturation field of the magnetization (H > H_s): this is a systematic offset due to our fitting procedure (which is designed to avoid any “false positive” THE detection) and hence reflects the maximum error in the magnitude of ρ^THE(H) shown in Fig. 1g.

We find that isolated skyrmion configurations (Fig. 1a–c) consistently generate a larger ρ^THE(H), despite their small n_sk(MFM). Contrary to expectations from Eq. (1), dense skyrmion arrays (Fig. 1d–f) with larger n_sk(MFM) typically display a smaller ρ^THE(H). The generic large ρ^THE(H) for isolated skyrmions can be further confirmed using the evolution of a dense skyrmion array with increasing H, as shown in Fig. 1d–f. Here, ρ^THE(H) rises as the skyrmion lattice dissolves into isolated skyrmions before reaching a FM phase with ρ^THE(H) ≈ 0. To further probe this discrepancy between transport and imaging experiments, we evaluate Δ_ρ/M ≡ n_sk(THE)/n_sk(MFM)^{9,11,12,18,19,25}, which quantifies the effective topological charge contributing to the measured THE, then compare this quantity for different skyrmion configurations imaged by MFM. As n_sk(MFM) increases, Δ_ρ/M is systematically reduced (Fig. 1g). Hence for a skyrmion lattice, we have n_sk(THE) ≈ n_sk(MFM) and for isolated skyrmions n_sk(THE) > n_sk(MFM). The strong influence of the skyrmion configuration on ρ^THE underlines the crucial role of κ in determining the magnitude of ρ^THE(H). We therefore examine the transition between isolated and dense arrays of skyrmions, and the influence of κ on the measured THE.

Figure 2a depicts the T − κ space and hence the wide range of spin textures, which we can experimentally access by varying T (5–300 K), the Fe/Co thickness ratio, and the number of repeats of [Fe/Co] in a multilayer stack. Using our MFM images, we identify the transition (blue shaded region in Fig. 2a) between isolated skyrmions and a dense, disordered skyrmion lattice⁵ and track its dependence on T and κ (see Section 1 of the Supplementary Information for the evolution of magnetic textures across T and κ). We explore the impact of this transition on the THE by correlating the evolution of ρ^THE across the T − κ phase space. Figure 2b displays ρ^THE(T, κ) curves for individual multilayers studied as a function of T, while Fig. 2c shows ρ^THE(κ) curves at a fixed T with varying Fe/Co. The location of the maximum in ρ^THE (Fig. 2b, c) tracks the phase boundary between isolated skyrmions and a disordered skyrmion lattice tuned by T, κ. This confirms that the degree of THE enhancement is closely linked to proximity to the phase boundary.

Fig. 2: Evolution of ρ^THE across the phase transition between isolated skyrmions and a disordered skyrmion lattice in T − κ space.

a Phase diagram of skyrmion configurations in T − κ space obtained by varying T and Fe/Co composition. The shaded area in blue is the transition region from isolated skyrmions to a disordered skyrmion lattice revealed by MFM imaging. The spatial evolution of skyrmion configuration is analyzed using Delaunay triangulation: details of the analysis including the nearest-neighbor (NN) coordination number and NN angular orientation can be found in Section 1 of the Supplementary Information⁵. Details of our estimation of the T-dependent κ are presented in Section 2 of the Supplementary Information. b Evolution of ρ^THE with (T, κ). Arrows with dotted lines point to the specific samples in (a) for which ρ^THE is studied as a function of T. c Evolution of ρ^THE with κ at a fixed T. Arrows with dotted lines point to the various samples in (a) for which ρ^THE(κ) is studied at a fixed T. d Schematic T − κ phase diagram illustrating the phase boundary T_c(κ) and the use of the effective temperature ${T}^{\prime}=T\cdot \kappa$ to describe the phase transition. The error bars on ρ^THE in (b) and (c) are ≤2 nΩ cm.

Here we note that the sign of THE remains unchanged across the T − κ phase diagram. However, R₀ changes from positive to negative with T due to multiband transport⁹. This crossover consistently follows the local maximum in ρ^THE and the transition from isolated skyrmions to the lattice phase (for details see Sects. 4–6 of the Supplementary Information). Such correlation between the local enhancement of ρ^THE, skyrmion configuration, and sign reversal of R₀ suggests that these factors are influenced by systematic changes in the occupancy of the electronic bands while varying T and the Fe/Co composition. To understand the mechanism responsible for the R₀ sign reversal, we employed a tight-binding model together with ab-initio calculations of the electronic band structure of our multilayer stack. When electron-like spin-up states and hole-like spin-down states are both present near the Fermi energy, a T induced change in the individual carrier densities can lead to a sign change of R₀ due to compensation of the normal Hall signal from carriers with opposite charge. However, due to their opposing spin alignment, the interaction of these carriers with B_eff from the skyrmions gives rise to a THE which maintains the same sign from T = 5–300 K. A complete quantitative analysis of the band structure is beyond the scope of this work. However, these experimental observations provide valuable insight into the links between thermodynamic stability and charge transport in chiral magnetic textures.

We now return to the enhancement of the THE. Within the critical region surrounding a second-order phase transition, fluctuations of the incipient order parameter (η) dominate the material response. Hence, chiral spin fluctuations at all length scales may develop at the skyrmion lattice-phase boundary⁷. The spin chirality contribution to the THE originates from a non-zero scalar triple product [S_i ⋅ (S_j × S_k)], where S_i, S_j, and S_k form a cluster consisting of three fluctuating spins. Such clusters can generate Hall signals via an unconventional skew-scattering mechanism, which can greatly exceed the contribution from a stable skyrmion configuration^15,16,17,26. The size of these clusters can be as small as several atomic spacings, which is much smaller than the skyrmion radius determined by the macroscopic magnetic interaction parameters in our films. Upon approaching the phase boundary, such fluctuation-induced clusters are expected to proliferate throughout the material, hence generating considerable topological charge^15,16,17. The correlation of these fluctuating spin clusters is associated with the formation/destruction of long-range order, i.e., a stable skyrmion lattice configuration. Within this scenario, the spin correlation length of the skyrmion lattice (which for an infinite thermodynamic system would diverge at the transition) is distinct from the short length-scale spin fluctuations responsible for a non-zero scalar spin chirality. A detailed microscopic understanding of this fluctuating spin chirality in relation to the phase transitions in chiral magnetic systems remains to be established theoretically²⁶. As we approach the transition, the influence of such fluctuations should be revealed by power-law behavior in material properties, including the spin susceptibility and hence THE²⁰. The maximum observable THE magnitude is expected to saturate at a value corresponding to the maximal chiral spin cluster density (imposed by the atomic spacing). Consequently, the power-law enhancement is truncated close to the transition and there is no THE singularity.

Our experimental observations reveal that the phase transition between isolated skyrmions and a disordered skyrmion lattice is sensitive to κ, T, and H. In the following, we consider an effective temperature ${T}^{\prime}$, which accounts for the role of varying T and κ in controlling this phase transition. Figure 2d illustrates how the critical temperature T_c and stability κ_c defining the transition vary with κ and T, respectively, in accordance with experiments (Fig. 2a). The phase boundary between isolated skyrmions and a skyrmion lattice may be considered a smooth function T_c(κ)^22,23,24. In the vicinity of a phase transition (κ = κ_c ~ 1, within the transition region, Fig. 2a) the boundary can be described as T_c ⋅ κ ≈ constant and hence,

$$\frac{{\mathrm{d}}\left({T}_{\mathrm{c}}\cdot \kappa \right)}{{\mathrm{d}}\kappa }=\frac{{\mathrm{d}}\left({\mathrm{constant}}\right)}{{\mathrm{d}}\kappa }\Rightarrow \kappa \frac{{\mathrm{d}}{T}_{\mathrm{c}}}{{\mathrm{d}}\kappa }+{T}_{\mathrm{c}}=0\Rightarrow \frac{{\mathrm{d}}{T}_{\mathrm{c}}}{{\mathrm{d}}\kappa }\approx -{T}_{\mathrm{c}},$$

which in turn suggests an effective temperature $T^{\prime} =T\cdot \kappa$ with a critical value $T^{\prime} ={T}_{\mathrm{c}}\cdot {\kappa }_{\mathrm{c}}$ defining the transition. One may also consider the phase boundary in terms of varying H, however, the interdependence of κ and H, and its variation with T remain unclear. We therefore adopt the simpler picture presented in Fig. 2d, which allows a straightforward interpretation of our experimental observations in Figs. 1 and 2a–c.

Figure 3a shows a power-law behavior in Δ_ρ/M surrounding a critical value of ${T}_{\mathrm{c}}^{\prime}\approx 110\pm 15$ K. This is consistent with the presence of a second-order phase transition driven by T and κ. According to Landau theory, the amplitude of η grows as a power-law on the low-symmetry side of any second-order transition (which in our case corresponds to ${T}^{\prime} \, > \, {T}_{\mathrm{c}}^{\prime}$):

$$\eta \propto | T^{\prime} -{T}_{\mathrm{c}}^{\prime}{| }^{\beta },$$

whereas fluctuations of η on both sides of the transition scale as:

$$\langle {({{\Delta }}\eta )}^{2}\rangle \propto \frac{1}{| T^{\prime} -{T}_{\mathrm{c}}^{\prime}{| }^{\gamma }}.$$

Extracting the critical exponents β, γ from our data requires the identification of η for the skyrmion lattice configuration. In the ordered phase of a chiral magnet, a helical wavevector that describes the spin rotation period emerges, capturing the symmetry broken by η at the transition. Discrete Fourier transforms of a skyrmion lattice yield a strong peak in the structure factor at k = l⁻¹, where $l\propto 1/\sqrt{{n}_{\mathrm{sk}}({\mathrm{MFM}})}$ is the skyrmion lattice parameter and k rises continuously from zero with the emergence of a lattice. We, therefore, expect η to scale with $\sqrt{{n}_{{\mathrm{sk}}}({\mathrm{MFM}})}$, allowing analysis of the transition using the conventional scaling approach

$${n}_{\mathrm{sk}}({\mathrm{MFM}})\sim | T^{\prime} -{T}_{{\mathrm{c}}}^{\prime}{| }^{2\beta },\qquad T^{\prime} \, > \, {T}_{{\mathrm{c}}}^{\prime}.$$

(3)

Fig. 3: Insights into the critical behavior captured by the topological Hall effect (THE).

a Power-law behavior of Δ_ρ/M ≡ n_sk(THE)/n_sk(MFM) with effective temperature ${T}^{\prime}=T\cdot \kappa$, and a critical point at ${T}_{\mathrm{c}}^{\prime}=110\pm 15\,{\mathrm{K}}$. Insets show the estimation of critical exponents, γ for ${T}^{\prime} \, < \, {T}_{\mathrm{c}}^{\prime}$ and γ + 2β for ${T}^{\prime} \, > \, {T}_{\mathrm{c}}^{\prime}$. b Evolution of n_sk(MFM) with ${T}^{\prime}$, showing a power-law behavior for ${T}^{\prime} \, > \, {T}_{\mathrm{c}}^{\prime}$ and the estimation of the critical exponent 2β. c Identification of the critical region T_c(κ) using three different approaches, namely, the local maximum in ρ^THE with T and κ, the power-law rise in Δ_ρ/M and the transition region (from isolated skyrmions to a disordered skyrmion lattice) indicated by MFM. The error bar in the estimated exponents reflects the error in the slope of the linear fits shown in the insets to (a and b). The same error bar is reflected through the shaded regions in the power-law fits. Error bars on ${\rho }_{\mathrm{max}}^{\mathrm{THE}}$ in (c) reflect the variation in the peak position of ρ^THE between Fig. 2b, c. The shaded region for T_c(κ) from Δ_ρ/M in (c) reflects a conservative estimate of ±35% variation in κ, resulting from an upper bound in the estimation of the exchange (A) and DMI(D) constants as well as K_eff from magnetization measurements⁵.

Fluctuations of η in the vicinity of the phase transition create a topological charge and hence contribute an effective skyrmion density n_fl ≈ 〈(Δη)²〉 (n_fl ≫ n_sk(MFM)) to the total measured n_sk(THE). Consequently,

$${{{\Delta }}}_{\rho /{\mathrm{M}}}=\frac{{n}_{{\mathrm{sk}}}({\mathrm{THE}})}{{n}_{{\mathrm{sk}}}({\mathrm{MFM}})}=\frac{{n}_{{\mathrm{sk}}}({\mathrm{MFM}})+{n}_{{\mathrm{fl}}}}{{n}_{{\mathrm{sk}}}({\mathrm{MFM}})}\approx \frac{{n}_{{\mathrm{fl}}}}{{n}_{{\mathrm{sk}}}({\mathrm{MFM}})}.$$

Therefore, we may model the power-law rise in Δ_ρ/M as follows:

$${{{\Delta }}}_{\rho /{\mathrm{M}}} \sim \frac{1}{| T^{\prime} -{T}_{{\mathrm{c}}}^{\prime}{| }^{\gamma }},\qquad T^{\prime} \, < \, {T}_{{\mathrm{c}}}^{\prime}$$

(4)

$${{{\Delta }}}_{\rho /{\mathrm{M}}} \sim \frac{1}{| T^{\prime} -{T}_{{\mathrm{c}}}^{\prime}{| }^{2\beta +\gamma }},\qquad T^{\prime} \, > \, {T}_{{\mathrm{c}}}^{\prime}$$

(5)

where the extra 2β in the exponent above ${T}_{{\mathrm{c}}}^{\prime}$ originates from the appearance of a stable skyrmion lattice as indicated in Eq. (3).

Fits to Eqs. (3) and (4) to extract γ and β are shown as red lines in Fig. 3a (for ${T}^{\prime} \, < \, {T}_{{\mathrm{c}}}^{\prime}$) and 3b, respectively, where we use two independent data sets to obtain exponents γ ≈ 1.88 ± 0.43, and β ≈ 0.31 ± 0.05. These values suggest a three-dimensional Heisenberg spin system in which the exponents are modified by competing spin interactions in a quasi-two-dimensional environment. This is in agreement with earlier studies of thin-film magnets²⁷, which consistently reveal an increased γ and a reduced β with respect to the three-dimensional Heisenberg values 1.39 and 0.36, respectively²⁸. To independently crosscheck our estimate of the exponents γ and β, we also fit the discrepancy in the skyrmion lattice regime using Eq. (5) (Fig. 3a for ${T}^{\prime} \,> \, {T}_{c}^{\prime}$). We obtain the combination (γ + 2β) = 2.61 ± 0.44, which is consistent with the γ and β extracted individually from our fluctuation analysis, adding credence to the validity of these critical exponents.

As discussed above, the finite size of the short length-scale fluctuating spin clusters imposes an upper limit on the magnitude of n_fl, which is expected to saturate as the system approaches ${T}_{{\mathrm{c}}}^{\prime}$. Experimentally, we can only modulate ${T}^{\prime}$ in discrete steps, so we cannot tune ${T}^{\prime}$ continuously through ${T}_{{\mathrm{c}}}^{\prime}$ to reveal the expected saturation in n_fl and hence THE. However, our experimental Δ_ρ/M clearly demonstrates the anticipated power-law scaling on either side of ${T}_{{\mathrm{c}}}^{\prime}$, systematically varying from 1–10³. Using Eq. (1) and the values of ∣R₀∣ ≈ 0.5–16 nΩ cm/T measured in our multilayer films, we estimate that Δ_ρ/M should saturate at a value 10⁴–10⁵ as $| {T}^{\prime}-{T}_{{\mathrm{c}}}^{\prime}| \to 0$. From the power-law trends shown in Fig. 3a, we deduce that this saturation is only visible for $| {T}^{\prime}-{T}_{{\mathrm{c}}}^{\prime}| \, < \, 20$ K, which lies beyond the minimum $| {T}^{\prime}-{T}_{{\mathrm{c}}}^{\prime}| \approx 40\pm 15$ K accessed during our experiments. It may therefore be possible to engineer a further THE increase of at least another order of magnitude, by tuning skyrmion-hosting multilayers more closely towards their isolated skyrmion/disordered lattice transition. Finally, we note that the presence of a finite population of skyrmions for ${T}^{\prime} \, < \,{T}_{{\mathrm{c}}}^{\prime}$ is consistent with the nucleation-type transition by which the skyrmion lattice is established^22,23,24. Some of these isolated skyrmions may also be stabilized by local variations in the magnetic interactions due to disorder in our films.

Discussion

Our results are summarized in Fig. 3c, which highlights the transition between isolated skyrmions and a disordered skyrmion lattice, determined by real-space imaging of the spin textures as well as magnetotransport. The critical parameter governing the transition T_c(κ) is identified by three separate methods: the maximum in ρ^THE, MFM imaging, and the fluctuation-induced rise in Δ_ρ/M at ${T}_{\mathrm{c}}^{\prime}=110\pm 15$ K. All three data sets display considerable overlap, indicating the active role of critical spin fluctuations in determining the magnitude of THE in chiral magnetic films. The ensuing discrepancy of up to three orders of magnitude between n_sk(THE) and n_sk(MFM) in the vicinity of the phase transition may account for the widely-varying magnitudes of ρ^THE values previously reported in technologically-relevant chiral magnetic films^{9,10,11,12,18,25,29}. Our material platform allows the THE to be tuned from n_sk(THE) ≈ n_sk(MFM), which is typical for skyrmion crystals of B20 compounds^6,7,8, to n_sk(THE) > n_sk(MFM) in dilute skyrmion configurations characteristic of interfacial systems^{9,10,11,12,18,25,29}. This acute sensitivity of ρ^THE to the magnetic skyrmion configuration indicates the crucial role of chiral spin fluctuations and opens a promising avenue towards controllable topological spintronics.

Methods

Film deposition

Thin-film multilayers consisting of Ta(30)/Pt(100)/[Ir(10)/Fe(x)/Co(y)/Pt(10)]_N/Pt(20) (numbers in the parentheses are layer thickness in Å and N refers to the number of times the Ir/Fe/Co/Pt stack is repeated, x, y are varied between 0 and 6 Å and 5–10 Å, respectively) were deposited on thermally oxidized Si wafers by dc magnetron sputtering at room temperature, using a Chiron ultra-high vacuum multi-source system. The optimal growth rates for individual layers are Ta: 0.55 Å/s, Pt: 0.47 Å/s, Ir: 0.12 Å/s, Fe: 0.13 Å/s, Co: 0.2 Å/s. The base vacuum of the sputter chamber is 1 × 10⁻⁸ Torr and an argon gas pressure of 1.5 × 10⁻³ Torr is maintained during sputtering.

Electrical transport

The magnetotransport measurements were performed using a custom-built variable-temperature insert (VTI) housed in a high-field magnet, complemented by a Quantum Design Physical Property Measurement System (PPMS). Current densities as low as 10⁴ A/m² at 33 Hz were used to avoid current-driven perturbation of spin textures. Detailed analysis for the extraction of the topological Hall resistivity (ρ^THE(H)) can be found in our earlier works^5,9. The non-zero value of ρ^THE at H > H_S serves as a conservative estimate of the error bar in the extracted ρ^THE(H), which is ≤2 nΩ cm. This includes the systematic errors involved in data analysis.

Magnetization

Magnetization measurements were performed in the range T = 5–300 K and a magnetic field of ±4 T, using superconducting quantum interference device (SQUID) magnetometry, in a Quantum Design Magnetic Property Measurement System (MPMS) (see Section 2 of the Supplementary Information). M(H) loops were recorded with the applied field in-plane (hard axis) and out of the film plane (easy axis). The saturation magnetization M_S and the difference in saturation fields H_S along the easy and hard axes were used to estimate the effective magnetic anisotropy K_eff^5,9.

Magnetic interactions

M_S and K_eff were determined from SQUID magnetometry measurements. The detailed methods for estimatingthe exchange stiffness A and DMI were reported in our previous work⁵. In this work, we use the scaling laws $\frac{A(T)}{A(T=5\,{\mathrm{K}})}={\left[\frac{{M}_{\mathrm{S}}(T)}{{M}_{\mathrm{S}}(T = 5\,{\mathrm{K}})}\right]}^{1.5}$ and $\frac{D(T)}{D(T=5\,{\mathrm{K}})}={\left[\frac{{M}_{\mathrm{S}}(T)}{{M}_{\mathrm{S}}(T = 5\,{\mathrm{K}})}\right]}^{1.5}$ involving the T-dependent saturation magnetization M_S(T) to estimate A(T), D(T), and the resulting κ(T). Details can be found in Section 2 of the Supplementary Information.

Magnetic force microscopy (M F M)

Room temperature MFM experiments were carried out using a Veeco Dimension 3100 Scanning Probe Microscope. The MFM tips used (Nanosensors SSS-MFMR) were ≈30 nm in diameter, with low coercivity (≈12 mT) and ultralow magnetic moment (≈80 emu/cc). Samples were initially saturated in the out of plane (OP) direction using fields up to H = −0.5 T. The measurements were performed in OP fields starting from H = 0 and incrementally approaching H = +H_S, with a typical tip height of 20 nm. The field evolution of the spin textures is presented in Section 1 of the Supplementary Information. Low T (5–200 K) MFM imaging is carried out using a cryogenic frequency-modulated MFM system⁹. We used two commercial probes by Team Nanotec, model ML3 (35–40 nm Co alloy coating), with f₀ ≈ 75 kHz and k₀ ≈ 1 N/m. The sample was first stabilized at a given temperature and then magnetized in the OP direction by applying H > H_S. After saturation, MFM images were acquired at various field values as H was swept from −H_S to +H_S. The details of the MFM image analysis are reported in our earlier reports^5,9,30.

Calibration of applied magnetic field

Extensive field calibrations were performed using a reference Hall sensor and a reference Palladium sample to minimize the field offsets between different experimental setups. Sensor details and the calibration data together with analysis for the extraction of ρ^THE(H) are reported in our earlier works^5,9. An additional field-calibration crosscheck is performed for our MFM images by estimating the magnetization from the image itself. Here the magnetization of the image is estimated as the ratio of the effective area of the imaged surface polarized along the applied field to the total area of the image. The magnetization of the MFM image is then located on M(H) data recorded by the SQUID magnetometer.