Sr2RuO4 stands out among the unconventional superconductors as one of the few materials with a chiral order parameter1,2. The tetragonal crystal structure allows five unitary representations for a p-wave pairing symmetry1,3. One of these is the chiral order parameter, of the form kx ± iky, which is strongly suggested by muon spin relaxation4 and high-resolution polar Kerr effect measurements5. Very recently, nuclear magnetic resonance experiments demonstrated that the d-vector is not parallel to the c-axis and suggested possible chiral d-wave states6,7. Such chiral states are attracting renewed attention due to the possibility of hosting Majorana bound states, which in turn are of interest for topological quantum computing8,9,10. A key property of the chiral state is its double degeneracy in the orbital degree of freedom, with important consequences such as the existence of superconducting domains of different chirality and a spontaneous edge current. The major problem plaguing our understanding of Sr2RuO411 is that, although the chiral state seems probable, domains or edge currents have not been observed directly. Indications for their existence, however, have been found in transport experiments, which utilise Ru inclusions to form proximity junctions between Sr2RuO4 and a conventional s-wave superconductor12,13. A complication in the physics of Sr2RuO4 is that breaking of the tetragonal crystal symmetry due to Ru inclusions or a uniaxial strain can induce a different superconducting state with an enhanced superconducting transition temperature Tc ≈ 3 K14,15. Recent experiments suggest that this so-called 3-K phase may exhibit a non-chiral state with a single-component order parameter16,17. In this paper, we refer to the multi-component phase with Tc of around 1.5 K, associated with the pure bulk limit, as the “intrinsic phase” and the possible single-component phase, characterised by Tc ≈ 3 K, as the “extrinsic phase”.

The vast majority of experiments in the past two decades have been limited to bulk crystals, typically hundreds of microns in dimension. This is partly due to the unavailability of superconducting Sr2RuO4 films. The chiral domains, however, are expected to be no more than a few microns in size5,13. Moreover, the time-dependent switching noise observed in transport measurements suggests the domains are mobile12,13. We note here that the role of chiral domains resulting in hysteretic behaviour has been discussed in the Bi-Ni bilayer system18. The arbitrary configuration of the domains introduces an element of uncertainty. On the other hand, the energy cost associated with a chiral-domain wall (ChDW) grows per area19. It has been recently discussed that mesoscopic samples made of chiral p-wave superconductors could host multichiral states20,21, where the two kx ± iky chiral components are divided into superconducting domains, separated by ChDWs. This makes mesoscopic structures a promising platform to verify and potentially control the domains. Another interesting aspect of a ChDW is that it can act as a Josephson junction19 due to the local suppression of the order parameter, as schematically shown in Fig. 1a.

Fig. 1: Superconducting domain wall and the Sr2RuO4 microring.
figure 1

a Schematic of a chiral-domain wall (ChDW). η and η+ represent the degenerate chiral states meeting at a ChDW. The colour wheels represent the orbital phase of the chiral components, which wind in opposite directions. The two chiral wavefunctions overlap over a finite length L. As they locally suppress each other, a Josephson junction is formed. b False-colour scanning electron microscope (SEM) image of Ring A. The blue represents the Sr2RuO4 crystal, and the yellow represents silver paint used for making electrical contact. Close-up images of c Ring A and d Ring B. In both rings the outer radius is around 0.55 μm while the inner radius of Ring B (rin = 0.3 μm) is slightly larger than that of Ring A rin = 0.23 ± 0.04 μm. Each ring is connected to four transport leads and is sculpted out of a single crystal (around 0.7 μm thick) by a Ga+ focused ion beam.

Here, we present results of transport measurements on mesoscopic rings of Sr2RuO4, prepared by focused ion beam (FIB) milling of single crystals. Homogeneous structures, characterised by a sharp transition at around the intrinsic Tc of 1.5 K, show distinct critical current oscillations—similar to that of the classical DC superconducting quantum interference device (SQUID), consisting of two artificially prepared Josephson junctions. Despite the absence of conventional weak links, the interference pattern appears over the full temperature range below Tc while maintaining its overall shape. In contrast, the SQUID oscillations are entirely absent in rings that are in the extrinsic phase. These systems behave as standard superconducting loops: they exhibit the conventional Little–Parks (LP)22 Tc oscillations, which can only be observed near the resistive transition21. We also present calculations on the possible chiral-domain configurations for a p-wave superconducting ring, using the Ginzburg–Landau (GL) formalism. Experiments and calculations together make a convincing case for the existence of ChDWs in the intrinsic phase of Sr2RuO4.


Basic transport properties of single-crystal microrings

Single crystals of Sr2RuO4 were grown with the floating zone method23 and structured into microrings using Ga-based FIB etching. Figure 1b–d shows scanning electron microscope (SEM) images of Rings A and B. The inner and outer radii of Ring A are rin ≈ 0.21 μm and rout = 0.55 μm, respectively. Similar dimensions are used in Ring B: rin = 0.3 μm and rout = 0.54 μm. Both crystals have a thickness of around 0.7 μm.

The temperature-dependent resistance R(T) of both rings (presented in Fig. 2a, b) shows sharp superconducting transitions similar to that of bulk Sr2RuO4. The apparent enhancement of the resistance just above Tc in Fig. 2b could be attributed to changes in the current path24. The high quality of the sample is also evident by their particularly high residual resistivity ratio; RRR = R(300 K)/R(3 K) = 238 for Ring A and RRR = 177 for Ring B. To demonstrate that FIB milling does not alter the intrinsic characteristics of Sr2RuO4, we compare the R(T) of Ring A with the one measured before milling the crystal in Supplementary Fig. 1, which shows that Tc and the overall transport properties remain unchanged under structuring. Figure 2c shows the typical current–voltage V(I) behaviour at different temperatures. For both rings, the V(I) measurements exhibit negligibly small hysteresis even at temperatures far below Tc.

Fig. 2: Basic properties of the Sr2RuO4 microrings.
figure 2

Resistance as a function of temperature R(T) for a Ring A and b Ring B. a presents data for various measurement currents, and b was measured using 10 μA. The insets show the R(T) over a wider temperature range. Both rings exhibit highly metallic behaviour with residual resistivity ratio of 238 for Ring A and 177 for Ring B. c Current–voltage characteristics V(I) of Ring A at various temperatures. The colours represent different voltage regions: V < −0.1 μV (green), −0.1 < V < 0.1 μV (blue), and 0.1 μV < V (red).

Insights from theoretical simulations

Before presenting the results of transport measurements under a magnetic field, we examine the expected chiral-domain configurations in our structure. This is accomplished by performing detailed time-dependent GL simulations, under the assumption of a chiral p-wave order parameter, for microrings with nanostructured transport leads (similar to the one used in our experiments). The simulations show that the ring can host a mono-chiral-domain or a multi-domain state, depending on the parameters \(\frac{{r_{{\mathrm{in}}}}}{{\xi (T)}}\) and \(\frac{{r_{{\mathrm{out}}}}}{{\xi (T)}}\), which correspond to the inner and outer radii of the ring, scaled by the temperature-dependent coherence length \(\xi (T) = \xi \left( {T = 0} \right)\frac{{\sqrt {1 - t^4} }}{{1 - t^2}}\)25,26, where \(t = \frac{T}{{T_{\mathrm{c}}}}\), with Tc ≈ 1.75 K for Ring A and Tc ≈ 1.3 K for Ring B (shown in Fig. 2a, b). Based on our critical field measurements, we estimate ξ(T = 0) ~ 66 nm, which is the same as the bulk value for Sr2RuO4. Figure 3a shows the simulated Cooper-pair density |Ψ|2 of Ring A far below Tc, obtained by setting \(\frac{{r_{{\mathrm{in}}}}}{{\xi (T)}} = 2.5\) and \(\frac{{r_{{\mathrm{out}}}}}{{\xi (T)}} = 6.8\) (corresponding to T\(\lesssim\) 0.5Tc in our measurements). This state contains two distinct chiral domains, separated by a pair of ChDW. Within the domain wall the order parameter is reduced to about half of its original amplitude in the banks on each side, resulting in the formation of two parallel Josephson weak links. While the suppressed order parameter is unfavourable in terms of the condensation energy, the formation of such ChDW is favoured by the second term of the free energy in Supplementary Eq. (S11). Since the order parameter is suppressed at the sample edge, the second term gains importance with reducing sample size and may further enhance an inhomogeneous order-parameter state with a ChDW. The ChDW region extends over a length of the order of ξ. As shown in Fig. 3b, the presence of a magnetic field along the ring axis makes the positions of the ChDWs shift away from the middle of the arms since one of the chiral components is favoured by the magnetic field. The ChDWs, however, remain in the arms of the ring due to the strong pinning by the restricted dimensions.

Fig. 3: Simulated least-energy-state configurations of chiral p-wave microrings with nanostructured transport leads.
figure 3

a State at a temperature much below Tc without magnetic field (specifically, T ≈ 0.78K, rin = 2.5ξ, and rout = 6.8ξ). b State at the same temperature as (a) but with axial magnetic field. c State at a temperature close to Tc in zero magnetic field (specifically, T ≈ 1.45K, rin = 1.3ξ, and rout = 3.6ξ). These states contain ChDWs that act as weak links of a SQUID. The colour maps represent the Cooper-pair density |Ψ|2. In the panels (ac), the inner rin and outer rout radii correspond to those of Rings A and B. The upper halves of the panels show the Cooper-pair density for each chiral component |η±| of the corresponding states, for which the overall order parameter is expressed by Ψ = η+ k+ + ηk. d \(\left( {\frac{{r_{{\mathrm{in}}}}}{{\xi (T)}},\frac{{r_{{\mathrm{out}}}}}{{\xi (T)}}} \right)\). Phase diagram in the absence of a magnetic field. For a wider ring arm, the “Meissner” state without ChDW is more stable; for a narrower ring arm, the extended ChDW state is expected. The dashed line shows the evolution of the least-energy states with increasing temperature according to the actual parameters of Ring A.

Figure 3c shows the calculated chiral-domain configuration for Ring B, which also applies to Ring A at temperatures near Tc. This is obtained by setting \(\frac{{r_{{\mathrm{in}}}}}{{\xi (T)}} = 1.3\) and \(\frac{{r_{{\mathrm{out}}}}}{{\xi (T)}} = 3.6\) (corresponding to T ≈ 1.45 K for Ring A). As the arms of the ring are now considerably narrower on the scale of ξ(T), the contribution of the edge regions dominates the configuration of the order parameter. As a consequence, it becomes energetically favourable for the two chiral components to coexist over the entire ring. This state also produces a pair of parallel weak links due to the suppression of the order parameter |Ψ|, which extend over the arms of the rings. Figure 3d presents a phase diagram of the lowest energy states, calculated for various \(( {\frac{{r_{{\mathrm{in}}}}}{{\xi (T)}},\frac{{r_{{\mathrm{out}}}}}{{\xi (T)}}})\). Amongst these, mono-domain “Meissner” state can be stabilised by increasing the rout/rin ratio. In this state, the arms of the ring are unable to provide effective pinning of ChDWs. This scenario is explored in Supplementary Note 4 and Supplementary Fig. 5, where a ring with relatively wide arms (Ring E) approaches the mono-domain state at low temperatures. The evolution of the equilibrium domain configuration as a function of temperature for Rings A and B are represented by the dashed lines in Fig. 3d. This suggests that the rings are in one of the domain states shown in Fig. 3a and c at all temperatures below Tc, except in a narrow range around 1–1.2 K, where additional domain walls could appear in Ring A. As a general finding, our GL calculations show that ChDWs could spontaneously emerge in our mesoscopic rings and behave as stable Josephson junctions over a broad temperature range, resulting in a DC SQUID of intrinsic origin. The change of chirality across such junctions and its influence on their transport characteristics remain open questions and are worthy of further studies. Note that the GL formalisms for chiral p-wave and chiral d-wave superconductors have analogous form, and the segregation of chiral domains as discussed above is applicable to both cases.

Critical current oscillations

We examined the supercurrent interference of the rings by measuring Ic at each magnetic field H. The results are presented in Fig. 4, where we observe the same behaviour in both Rings A and B. Figure 4a, b shows the Ic of Ring A, measured for positive (Ic+) and at negative (Ic−) bias currents, taken at temperatures deep inside the superconducting state and close to Tc, respectively. For both temperatures, we observe distinct critical current oscillations, with the period corresponding to the fluxoid quantisation over the ring area. This interference pattern corresponds to that of a DC SQUID with a pair of parallel Josephson junctions. The junctions would also need to be symmetric each other; an imbalance in Ic could not produce the cusp-shaped minima of the patterns. The figure also shows −Ic−(−H) overlaid on its time-reversed counterpart, +Ic+(H). Figure 4c shows that the same SQUID oscillations appear in Ring B, only with a slightly smaller period (consistent with its slightly larger inner radius). The oscillations emerge spontaneously at the onset of superconductivity and continue down to TTc. More importantly, we find that the patterns are not distorted, despite the substantial variations in Ic(T) and ξ(T).

Fig. 4: SQUID oscillations observed in the Sr2RuO4 microrings.
figure 4

a Critical current as a function of magnetic field Ic(H) of Ring A measured at 0.78 and 1.50 K. The open blue circles show time-reversed critical current. c Ic oscillations in Ring B over a wide range of temperatures. The Ic values were obtained from I(V) measurements at each magnetic field. The rings were heated up to above 5 K between each magnetic field.

It is worth noting that, unlike the polar Kerr experiments, we find field cooling and zero-field cooling of the samples to yield the same results in our measurements. This, however, is to be expected in mesoscopic structures, where domain walls are strongly pinned to the confined regions in order to lower the free energy of the system (see Supplementary Note 2 and Supplementary Fig. 2 for more details). Such pinning mechanism is absent in the polar Kerr experiments, which are performed on bulk crystals5.

To demonstrate the robustness of the SQUID behaviour further, in Fig. 5a, b, we plot the magnetoresistance of Ring A, produced by the Ic oscillations over a wide range of temperatures. These are measured by applying a constant DC current ±I while sweeping the magnetic field H along the ring axis. Here, the resistance R is defined by the average of two voltages before and after current reversal at each magnetic field; R = [V(I) − V(−I)]/2I. When the measurement current exceeds the critical current Ic(H), the system is driven out of the zero-voltage regime of the V(I) and produces a finite resistance. Combining the results of a wide range of temperatures, Fig. 5a, b reveals that the SQUID oscillations emerge together with Ic at the onset of the superconducting transition. In Fig. 5c, d, we describe the shape of R(H), where in some cases the peaks can appear to be split or broadened. This is clearly due to a slight difference in the values of I, which causes the voltage peaks for ±I to appear asymmetrically. We observed a similar asymmetry in rings showing the LP effect27.

Fig. 5: Magnetotransport of Ring A.
figure 5

Magnetoresistance R(H) with measurement current 50 μA for a temperatures between 1.75 and 1.10 K and b temperatures between 0.94 and 0.60 K. c R(H) at 0.78 K. The Ic oscillations at this temperature are displayed in Fig. 4a. d The corresponding magnetovoltage when the measurement current is applied to one direction V+ and to the other direction V. The peaks (dips) in V+ (V) appear at different field values and hence double peaks appear in the resistance.

The magnetovoltage and field-dependent V(I) measurements are crucial in resolving an outstanding issue regarding previous reports of unconventional behaviour of Sr2RuO4 rings. Cai et al. have consistently observed magnetoresistance oscillations with unexpectedly large amplitude28,29, very similar to the data presented in Fig. 5. The reported magnetoresistance oscillations are also stable over a wide range of temperatures and, in some cases, show small dips around Φ0/2. As Fig. 5 demonstrates, however, the averaged resistance R could produce a very similar effect even when there is no splitting of the peaks in the raw magnetovoltage signal.

T c oscillations in rings with an extrinsic phase

We already mentioned that ChDWs can produce the observed Ic(H) oscillations by acting as Josephson junctions. This should be contrasted with the fluxoid-periodic behaviour of structures with a partial or full extrinsic phase, characterised by a noticeably broader transition which begins near 3 K (see Fig. 6e and Supplementary Fig. 4). We recently reported observations of the LP oscillations in such Sr2RuO4 microrings27, and here we demonstrate that those are of a fundamentally different nature than the Ic oscillations discussed in this report. For this, we compare the data from Ring A with those of Ring C (sample B in ref. 27), where the transition is considerably broader (Fig. 6e). This ring was prepared from a 2-μm-thick crystal with a Tc of 1.5 K. After microstructuring, however, the ring was found to have a higher Tc, with its transition already starting at 2.7 K. The magnetotransport measurements reveal that the ring itself is predominantly in the extrinsic phase, introduced by microstructuring (most likely due to a strain induced by FIB milling of the thick crystal). Compared to Rings A (RRR = 238) and B (RRR = 177), this structure has a smaller residual resistivity ratio RRR = 129. Nevertheless, the value of RRR is still substantial, indicating strong metallicity for Ring C. Figure 6a, b shows R(H) for temperatures within the resistive-transitions of Rings A and C (taken 1.67 K and 2.3 K, respectively). In both cases we find fluxoid-periodic oscillations, which we compare with simulated LP oscillations (the red curves).

Fig. 6: Comparison with a ring that exhibits the Little–Parks (LP) oscillations.
figure 6

a Magnetoresistance R(H) of Ring A near the transition temperature. b R(H) of Ring C. The amplitude of the observed oscillations (blue and green) in Ring A is much greater than the expectation for the LP oscillations (red). In contrast, the observed oscillations in Ring C are in good agreement with a simulation for the LP oscillations. Colour maps of the differential resistance dV/dI of c Ring A and d Ring C as functions of magnetic field and measurement current. The bright part corresponds to the critical current. The critical current of Ring C does not show any oscillation. e Resistance as a function of temperature of Ring C. The onset Tc is 2.7 K, and hence the ring is in the extrinsic phase, which we consider as a non-chiral state. Data for panels (b) and (e) are adopted from ref. 27.

The change of the transition temperature due to the LP oscillations is given by30:

$$\begin{array}{*{20}{c}} {\frac{{T_{\mathrm{c}}\left( H \right) \;-\; T_{\mathrm{c}}\left( 0 \right)}}{{T_{\mathrm{c}}\left( 0 \right)}} = - \left( {\frac{{\pi \xi \left( 0 \right)w\mu _0H}}{{\sqrt 3 {\mathrm{\Phi }}_0}}} \right)^2 \;-\; \frac{{\xi ^2\left( 0 \right)}}{{r_{{\mathrm{in}}}r_{{\mathrm{out}}}}}\left( {n - \frac{{\pi \mu _0Hr_{{\mathrm{in}}}r_{{\mathrm{out}}}}}{{{\mathrm{\Phi }}_0}}} \right)^2,} \end{array}$$

where Φ0 = h/2e is the flux quantum with the Planck constant h and the elementary charge e, and w = rout − rin is the width of a ring arm. The first term represents the effect of the Meissner shielding, and the second term corresponds to fluxoid quantisation. To convert the change of the transition temperature to the resistance variation, we assume that the R(T) curve does not change its shape under magnetic field and shifts horizontally by ΔTc(H) = Tc(H) − Tc(0). For the simulations in Fig. 6a, b, we used ξ(0) = 66 nm, 2rin = 0.55 μm, 2rout = 1.1 μm for Ring A, and 2rin = 0.7 μm, 2rout = 1.0 μm for Ring C. Both the period and amplitude of the oscillations for Ring C agree with those of the simulation. We therefore consider these to be the LP oscillations, driven by variations in Tc. For Ring A, however, the oscillation amplitude is substantially larger than what Tc variations can produce. Such large-amplitude magnetoresistance is driven by the Ic(H) oscillations instead. In Fig. 6c, d, we compare the Ic(H) of both rings at lower temperatures. In contrast to Rings A and B, the SQUID oscillations are completely absent in Ring C. Instead, for all temperature below Tc, we only observe a monotonous decay of Ic(H). We find the lack of Josephson junctions to be a common characteristic among structures with a dominant extrinsic phase. A further example of this is given in Ring D (Supplementary Note 3 and Supplementary Fig. 4).


Before adopting ChDW scenario as the origin of the observed Ic oscillations, we consider other known mechanisms for Ic oscillations. Firstly, even in a homogeneous loop SQUID-like behaviour may emerge depending on the size of the ring with respective to either the penetration depth λ or the coherence length ξ. Ic can be modulated by the circulating persistent current Ip, which varies linearly with the flux, and switches its direction at every increment of Φ0/2. This mostly results in a sawtooth-like modulation of Ic31, which cannot account for non-linear form of the patterns shown in Fig. 4. Furthermore, the magnitude of Ip is inversely proportional to the kinetic inductance LK, which depends on the penetration depth LKλ2(T). If the Ic oscillations were driven by circulating currents, their amplitude ΔIc would grow larger by lowering the temperature since ΔIcIp 1/λ2(T)31. This is clearly not the case for the Sr2RuO4 rings, where oscillation amplitude is unaffected by temperature (e.g. ΔIc ≈ 12 μA at both temperatures shown in Fig. 4a). SQUID oscillations can also emerge in loops without weak link, if the dimensions are much smaller than ξ(T) and λ(T)32. However, this is not applicable to our structures, where the radii and the width of the arms are several times larger than the characteristic length scales for TTc (e.g. for Ring A, ξ(T) ~ 0.07 μm and λ(T) ~ 0.19 μm at T = 0.78 K).

Secondly, Cai et al. attributed the large-amplitude magnetoresistance of their Sr2RuO4 rings to current-excited moving vortices28,29. As demonstrated by Berdiyorov et al.33, this mechanism can only produce large-amplitude oscillations over a finite temperature range, typically down to T ~ 0.95Tc (e.g. see Fig. 6b of ref. 33 and Fig. 2 of ref. 34). This is not the case for the Sr2RuO4 rings that are in the intrinsic (1.5-K) phase, as the magnetoresistance oscillations appear for all T < Tc (see Fig. 5a, b and Fig. 3a in ref. 28).

Thirdly, geometrical constrictions (e.g. bridges and nanowires) can serve as Josephson junctions, as long as their dimensions are comparable to ξ. The current-phase relation (CPR) of such junctions is defined by the ratio of ξ(T) to the length of the weak link L. Since ξ(T) varies with temperature while L remains fixed, the CPR of such weak links is strongly temperature dependent. Generally, lowering the temperature transforms the CPR from sinusoidal to a sawtooth-like function, which ultimately turns into multivalued relations once L ≥ 3.5ξ(T), corresponding to the nucleation of phase-slip centres35,36,37. The multivalued CPR manifests itself as a hysteretic V(I) relation, which is a well-known characteristic of constriction junctions at TTc38,39. This is in direct contrast to the V(I) curves of the Sr2RuO4 rings, which show negligible hysteresis for temperatures as low as 0.2Tc (see Fig. 2c). Furthermore, the interference patterns taken at over wide range of temperatures show the same overall shape, with characteristically round lobes (Fig. 4). This could not be produced by constriction junctions, as the interference pattern would be heavily deformed by the pronounced changes in ξ(T)/L with temperature. In case of ChDWs, however, the length of the junction barrier is determined by the coherence length and therefore has a temperature dependence similar to ξ(T). Hence, a ChDW junction can maintain a relatively fixed ξ(T)/L(T) ratio for different temperatures. This would agree with the lack of hysteresis in our V(I) measurements (Fig. 2c) and the unperturbed shape of the interference patterns (Fig. 4).

Lastly, we exclude the possibility of forming accidental proximity junctions by Ru inclusions or any other normal metal within the Sr2RuO4 crystal. Apart from their absence in the SEM images taken while the milling of the rings, inclusions would induce an extrinsic 3-K phase. The crystals, however, show no such enhancement of Tc either before or after FIB processing. Moreover, the (single) sharp resistive transitions of Rings A and B could not be produced in the presence of normal metal weak links. Accidental tunnel junctions, formed by nanocracks or grain boundaries, can also be excluded due to the high metallicity of our samples. In summary, the Josephson effect found in Sr2RuO4 microrings cannot be attributed to conventional types of weak link such as constriction junctions, kinematic vortices (phase-slip lines), proximity and tunnel junctions.

To summarise, our simulations of a chiral p-wave order parameter show that a mesoscopic loop with nanostructured transport leads can host a multi-domain state. The degenerate chiral states are separated by ChDWs located in the arms of the ring, where a pair of parallel Josephson junctions is formed due to the local suppression of both chiral states. We examined the existence of such junctions by performing transport experiments on Sr2RuO4 microrings. The rings with a sharp transition near 1.5 K show distinct Ic oscillations, similar to that of a DC SQUID with a pair of Josephson junctions with matching Ic. The junctions emerge together with the superconducting transition and are present for all temperatures below Tc. In contrast, for Sr2RuO4 rings with an extrinsic (3-K) phase, the Josephson junctions are entirely absent. Such rings show standard Little–Parks oscillations near Tc, which can be properly modelled, but no critical current oscillations. Our findings suggest that the Josephson junctions are an inherent property of the order parameter, and make a compelling case for the existence of ChDWs in the intrinsic (1.5-K) phase of Sr2RuO4. We should note that our present results formally do not distinguish the type of degenerate states responsible for the formation of the junctions; our transport measurements would also be consistent with domain walls of helical states, as well as of spin-singlet chiral states. This work also demonstrates that the combination order parameter simulations with mesoscopic structures can be instrumental in the study of superconducting domains and will, in coming experiments, allow for detailed design and understanding of a system before the actual fabrication.


Microring fabrication

Sr2RuO4 single crystals were prepared with the floating zone method23, and their transition temperature Tc before the sample fabrication was confirmed to be 1.50 K using a compact AC susceptometer40 in a Quantum Design PPMS. We crush the crystal into small pieces to obtain thin crystals with the thickness of approximately 1 μm. Although Sr2RuO4 is chemically stable in the ambient condition, we find that small crystals can degrade in the air. Therefore, freshly crushed crystals were used. The crystal is placed on a SrTiO3 substrate, where it is contacted by either gold or silver for transport measurements. For Rings A, C and D, two pads of high-temperature-cure silver paint (6838, Dupont) are attached to the two sides of the crystal. The paint is then cured at 500 °C for 20 min. In case of Rings B and E however, the crystals are contacted using a combination of electron-beam lithography and sputter deposition of gold. Once a crystal is contacted by the gold or silver paint, a 100-nm-thick layer of SiO2 is deposited using electron beam evaporation to protect the crystal during structuring. The contacts and the crystal underneath are then cut with a Gallium FIB to produce a four-wire arrangement. Lastly, the microrings are structured using the FIB (30 kV, 20 pA).


Transport measurements were performed in a 3He refrigerator (Heliox, Oxford Instruments) down to 0.3 K. In the DC resistance measurement, we flip the direction of the measurement current to subtract the contribution of the thermoelectric voltage, and the resistance R is defined to be R = [V(I) − V(−I)]/2I. The transition temperature shift due to the LP oscillations is calculated to be approximately 10 mK by using Eq. (1). Therefore, temperature stability during the magnetoresistance measurement must be much smaller than this value. By putting a 80-Ω by-pass resister in parallel to the heater and by tuning the PID values of the temperature controller, we achieved a temperature stability of 100 μK. Current–voltage V(I) measurements are performed under constant temperature and magnetic field with triangular current waves of frequency 2 mHz.


For details of the Ginzburg–Landau simulations, we refer to the formalism of ref. 20, and the additional discussion in the Supplementary information.